Electromagnetic Instabilities in an Inhomogeneous Plasma presents a comprehensive survey of the theory of electromagneti
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Pages [319] Year 1992
Table of contents :
General dispersion relation for a collisionless inhomogeneous plasma. Simplified dispersion relations. Inertial magnetodrift instabilities in a collisionless finiteB plasma. Alfven magnetodrift instabilities in a finiteB plasma. Lowfrequency instabilities in a twoenergy component plasma. Electromagnetic instabilities in a strongly inhomogeneous plasma. Description of perturbations in a plasma with transverse current. Transversecurrentdrive instabilities. Hydrodynamic description of collisional finiteB plasma perturbations. Starting equations for basic types of hydrodynamic perturbations. Instabilities in a collisional finiteB plasma. Instabilities in forcefree and almost forcefree highB plasmas. Electromagnetic KelvinHelmholtz instabilities. Electromagnetic drift instabilities in a lowB plasma. Allowance for magneticfield curvature effects. Magneticfield curvature effects in kinetic and twofluid instabilities. References. Index.
E le c tr o m a g n e tic In sta b ilitie s in an In h o m o g en eo u s P la sm a
Plasm a Physics Series Series Editor: P ro fe s s o r E W L ain g , University of Glasgow
A d v is o ry P a n e l D r J L ac in a, Czechoslovak Academy of Sciences P ro fe s s o r M A H e llb e rg , University of Natal P ro fe s s o r A B M ik h a ilo v sk ii, I V Kurchatov Institute of Atomic Energy P ro fe s s o r K M iy a m o to , University of Tokyo P ro fe s s o r H W ilh e lm sso n , Chalmers University of Technology
O t h e r books in the s e r i e s
A n I n tr o d u c tio n to A lfv e n W aves R Cross M H D a n d M ic ro in s ta b ilitie s in C o n fin ed P la s m a W M Manheimer and C N LashmoreDavies T ra n s itio n R a d ia tio n a n d T ra n s itio n S c a tte rin g V L Ginzburg and V N Tsytovich R a d io fre q u e n c y H e a tin g o f P la s m a s R A Cairns P la s m a P h y s ic s v ia C o m p u te r S im u la tio n C K Birdsall and A B Langdon
Plasm a Physics Series
Electromagnetic Instabilities Inhomogeneous Plasma
A B Mikhailovskii I V Kurchatov Institute of Atomic Energy Moscow
Translation Editor: P ro fesso r E W L aing
In stitu te of Physics Publishing Bristol, Philadelphia and New York
in an
© IO P Publishing Ltd 1992 All rights in editions in a Soviet language are retained by the author. All rights reserved. No part of this publication may be reproduced, stored in a retrieval system or transm itted in any form or by any means, electronic, mechanical, photocopying, recording or otherwise, without the prior permis sion of the publisher. Multiple copying is perm itted in accordance with the term s of licenses issued by the Copyright Licensing Agency under the terms of its agreement between the Committee of ViceChancellors and Principals.
British Library Cataloguing in Publication Data Mikhailovskii, A. B. (Anatolii Borisovich) Electromagnetic instabilities in an inhomogeneous plasma.  (Plasm a physics series) I. Title II. Series 530.401 ISBN 0750301821
Library o f Congress CataloginginPublication Data are available
Published by IO P Publishing Ltd, a company wholly owned by the Institute of Physics, London, by IOP Publishing Ltd Techno House, Redcliffe Way, Bristol BS1 6NX, England 335 East 45th Street, New York, NY 100173483, USA US Editorial Office: IO P Publishing Inc., The Public Ledger Buildings, Suite 1035, Independence Square, Philadelphia, PA 19106, USA Typeset in TftX at IO P Publishing Ltd P rinted in G reat B ritain by Galliard (Printers) Ltd, Great Yarmouth, Norfolk
Preface
The aim of this book is a comprehensive presentation of the the ory of electromagnetic instabilities in a magnetized inhomogeneous plasma, mainly in the classical approximation of straight and par allel magnetic field lines. In addition, the effects of magnetic field curvature are taken into account. The electromagnetic instabilities are more significant than the wellknown electrostatic ones when plasma pressure is finite or large compared with magnetic field pressure. Such plasmas are called finite or high/? plasmas. The presentation is mainly given from the general physics view point. Therefore, the book is useful for everyone interested in the current development of plasma theory. Confinement of finite or high/? plasmas is one of the main ob jectives in laboratory controlled fusion investigations. At the same time, such plasmas are widely representative of the state of matter in space. Therefore, the book will facilitate the labour of experts on controlled fusion and space physics who are engaged in electromag netic instabilities. Keeping in mind this category of readers, in the course of presentation the author mentions the specific problems of laboratory and space physics which led to the notions of instabilities considered. The study of the instabilities is based on analytical approaches. These approaches are presented in the book in detail. This allows one to show the main physical regularities in the types of instabilities discussed with minimal efforts. Then the reader can reproduce all the results given in the book, and, if necessary, make his/her own contribution in this new branch of plasma theory. A considerable part of the results given in the book has been obtained by the author and his colleagues. The same concerns the methods and the problem statements. At the same time, the main results in the fields which this book covers which have been obtained v
VI
Electromagnetic Instabilities in an Inhomogeneous Plasma
by other authors, both Soviet and Western, are also presented. The book contains reviews, including critical remarks, of publica tions on every type of instabilities, that should help the reader in orientation on the literature. I am indebted to Drs V A Pilipenko and 0 G Onishchenko and my daughter Lyudmila A Mikhailovskaya for their assistance during the preparation of the manuscript. A B Mikhailovskii February 1992
Contents Preface Introduction 1 1.1 1.1.1 1.1.2 1.1.3 1.1.4 1 .2
1.2.1 1.2.2 1.2.3 1.2.4 1.3 1.4 2 2 .1
2.1.1 2.1.2 2.1.3 2.1.4 2.1.5
General D ispersion R elation for a Collisionless Inhom ogeneous Plasm a General form of the dispersion relation The dispersion relation for a homogeneous plasma Potentials of the perturbed field Electrodynamic equations in terms of4>,A2,A 3 General form of the dispersion relation for an inhomogeneous plasma Equilibrium Constants of motion Equilibrium distribution function Equation of transverse equilibrium Equilibrium trajectories Perturbed distribution function Permittivity of an inhomogeneous plasma Simplified D ispersion R elations Permittivity and dispersion relations for lowfrequency longwavelength perturbations Permittivity for longwavelength lowfrequency perturbations Splitting of the general dispersion relation in the approximation k±p — 0 Dispersion relations for lowfrequency longwavelength perturbations with kz = 0 Permittivity of a forcefree plasma Permittivity of a plasma flow with an inhomogeneous velocity profile
V
xvii 1
3 3 4 5 6 8 8 9 10 11 12
14 16 17 18 19 20 22 23 vii
viii
Electromagnetic Instabilities in an Inhomogeneous Plasma
2.1.6
Allowance for the interaction of Alfven perturbations with resonant particles Dispersion relations for shortwavelength perturbations with kz = 0 Perturbations with E _L Perturbations with E  B 0 Oblique shortwavelength perturbations Lowfrequency perturbations Perturbations with u ~ Oi Electromagnetic perturbations in a low/? plasma
2.2 2.2.1 2.2.2 2.3 2.3.1 2.3.2 2.4 3 3.1 3.2 3.3 3.3.1 3.3.2 3.3.3 3.3.4 3.4 3.4.1 3.4.2 3.5 4 4.1 4.1.1 4.1.2 4.2 4.2.1 4.2.2 4.2.3 4.3 4.3.1
Inertialess M agnetodrift Instabilities in a C ollisionless Finite/? Plasm a Tserkovnikov’s magnetodrift instabilities Transverse drift waves Instabilities with finite kz Electron drift wave branch Magnetodrift oscillation branch Influence of finite /? on the ion temperaturegradient drift instability Suppression of the ion temperaturegradient drift instability in a plasma with /? > 1 Instabilities with large kz Influence of finite /? on the ion temperaturegradient drift instability Suppression of the ion temperaturegradient drift instability in a plasma with /? >> 1 Systematization of the results A lfven M agnetodrift Instabilities in a Finite/? Plasm a Alfven perturbation branches for finite /? and k±p\ —►0 Perturbations with kz = 0 Perturbations with kz ^ 0 Interactions of driftAlfven waves with resonant particles for kz = 0 Resonant interaction of the waves with electrons Resonant interaction of waves with ions The zeroth approximation in the parameters u>/kzVTe and kzvxi/to Role of resonant particles in Alfven perturbations with kz vt\ < u < kzvTe The zeroth approximation in the parameters u / k zVTe and kzVT\/w
24 25 26 30 31 81 32 33 35 38 40 42 43 45 48 49 47 48 49 51 53 55 55 56 57 57 58 59 59 59
43.2 43.3 4.4 4.5 5 5.1 5.2 5.3 5.3.1 5.3.2 6
Contents
ix
Interaction of waves with resonant electrons Ion resonance of the type u = kzvz + u di for u;/u>di < 0 Role of resonant particles in Alfven perturbations with u ~ kzvTl for VT = 0 Correspondence between Alfven perturbations and the flute instability in finite/? plasma
60 61
Lowfrequency Instabilities in a Twoenergy C om ponent Plasm a Gradient excitation of the fast magnetoacoustic waves by highenergy particles Driftmirror instability Gradient excitation of Alfven waves by highenergy particles Instability of longwavelength Alfven perturbations Instability of the shortwavelength Alfven perturbations
63 65 67 71 72 75 76 77
E lectrom agnetic Instabilities in a Strongly 80 Inhom ogeneous Plasm a Shortwavelength perturbations in a finite/? Maxwellian 6.1 84 plasma 85 6.1.1 The k±pe —►0 approximation 86 6.1.2 Electron damping of the iondrift oscillation branch 88 6.1.3 Plasma with cold electrons 88 Driftcyclotron instability in a finite/? plasma 6.2 90 6.2.1 Electrostatic magnetodrift driftcyclotron instability 91 6.2.2 Electromagnetic driftcyclotron instability Electromagnetic lowerhybriddrift (highfrequency 6.3 93 drift) instability in a finite/? plasma 6.3.1 Coldelectron plasma 94 6.3.2 Plasma with finite electron temperature 94 94 Drift losscone instability in a finite/? plasma 6.4 95 6.41 Coldelectron plasma 96 6.4.2 Hotelectron plasma 97 6.4.3 Stabilization by a warm plasma 98 Ioncyclotron excitation of transverse drift waves 6.5 99 6.5.1 Ioncyclotron instability of hydrodynamic type 99 6.5.2 Kinetic magnetodrift ioncyclotron instability Shortwavelength driftAlfven ( s d a ) perturbations in a 6.6 100 finite/? plasma 100 6.6.1 Oscillation branches 101 6.6.2 Hydrodynamic treatment of the s d a perturbations 6.6.3 Interaction of the s d a perturbations with resonant 101 electrons
X
Electromagnetic Instabilities in an Inhomogeneous Plasma
6.7
Ioncyclotron instability of the
7
D escription of Perturbations in a Plasm a with Transverse Current
7.1
sda
perturbations
102 104
7.8
Electron equilibrium Ion equilibrium Ion distribution just after applying an alternating electromagnetic field to a plasma Distribution of partially relaxing ions Ions in a rotating plasma Magneticfield gradient Equilibria previously discussed in the literature Modification of the permittivity tensor General remarks The highfrequency approximation for £qo Dispersion relation for electrostatic perturbations The lowfrequency approximation, with respect to electrons, for (3 ('ti Electromagnetic perturbations with u '±i ' t \ Allowance for the gradient effects Instabilities with kj_pe > 1 Plasma with f3 > 1 Oblique waves in a plasma with 3 —r 0 Electroncyclotron harmonics Instabilities in a rotating plasma
1
and 132 132 133 134 134
134 135 136 136
137
H ydrodynam ic D escription o f Collisional 139 Finite/? Plasm a Perturbations 142 Drift hydrodynamics of a finite/? plasma Multimoment transport equation set 143 Reducing the multimoment transport equation set by expansion in 1 /5 147 The simplest hydrodynamic equation set allowing for 149 dissipative effects of order u 11>\ Agreement between the moment method and the drift 151 kinetic equation method Hydrodynamic description of inertial perturbations, 152 neglecting dissipative effects Allowance for the transverse heat conductivity and 155 transverse viscosity in perturbations with d /d z = 0 Starting Equations for Basic Types o f H ydrodynam ic Perturbations Dispersion relations for perturbations with kz = 0, neglecting dissipative effects Inertialess perturbations Inertial perturbations Dispersion relations for perturbations with kz = 0 allowing for dissipative effects Effects of order uj/ v^ Effects of order v\k\p1 / u Allowance for the heat transfer Dispersion relation for transverse drift waves Starting equations for oblique perturbations Perturbations with kz ~ ( ^ u ) 1/2/ ^ Perturbations with kz ~ u>/vTl in a weakly collisional plasma Perturbations in a highf3 plasma with kz > lj/ vt\
156 158 158 160
161 161 163
164 165
166 166 167 169
xii
Electromagnetic Instabilities in an Inhomogeneous Plasma
10.3.4 Perturbations in a strongly collisional plasma 10.4 Dispersion relations for inertialess perturbations in a plasma flow with an inhomogeneous velocity profile 10.41 Starting equations 10.4*2 Inertialess perturbations in a plasma flow with V n0 = V T0 = 0 10.43 Allowance for the density gradient 10.5 Description of Alfven perturbations in a plasma flow with an inhomogeneous velocity profile 10.5.1 Alfven perturbations when neglecting drift effects 10.5.2 Magnetic viscosity tensor for dU/dx ^ 0, with Bx ^ 0 10.5.3 Allowance for the drift effects for Vn 0 = VT0 = 0 10.5.4 Dispersion relation of smallscale Alfven perturbations for V n 0 = VT 0 = 0 10.5.5 Dispersion relation of smallscale Alfven perturbations for V n 0 ^ 0, with VT 0 = 0 11 1 1 .1
11.1.1 11.1.2 11.1.3 11.1.4 1 1 .2
11.2.1 11.2.2 11.3 11.4 11.5 12 1 2 .1
12.1.1 12.1.2 12.1.3 1 2 .2
12.2.1 12.2.2
Instabilities in a Collisional Finite/? Plasm a Magnetodrift entropy instabilities Entropy perturbations when neglecting dissipative effects The role of dissipative effects of order u>/v\ The role of dissipative effects of order (k^prfvJu The role of the heat transfer Inertialess instabilities with kz ~ (ve^)1/2/vxe Magnetodrift instability in a plasma with VT = 0 Instability of electron drift waves in a plasma with VT ^ 0 Inertialess instabilities with kz ~ u>/vTl in a weakly collisional plasma Inertialess instabilities in a strongly collisional plasma Inertial instabilities Instabilities in Forcefree and Alm ost Forcefree High/? Plasm as Inertialess instabilities in a forcefree collisionless high/? plasma Plasma with Te = T\ and j3 —> 00 Plasma with Te j= TJ and /? —►00 Instability boundaries for finite j3 Inertialess instabilities in an almost forcefree collisionless high/? plasma Perturbations with small kz Perturbations with Vt\ u>/vTe are derived in sections 2 . 2 and 2.3. Finally, in section 2.4 the case of a low(3 plasma is considered. Note that one of the important results of this chapter is the splitting of the general dispersion relation for lowfrequency longwavelength perturbations into two mutually independent dispersion relations. One of them corresponds to the Alfven or inertial perturbations, while the second corresponds to magnetoacoustic and some other perturbations called, in general, inertialess perturbations. Such a splitting for arbitrary angles between k and B 0 was revealed in [1.8], and that splitting for the case k • B 0 = 0 was shown in [1.6]. This considerably facilitates the analysis of the lowfrequency longwavelength perturbations. In addition, as shown in [2 . 1 , 2 .2 ], the splitting is revealed under the hydrodynamic description of pertur bations (see also Chapters 9 and 10). Moreover, according to [2.3] (see also [2.4]), the same holds in the case of largescale perturba tions which cannot be described by the WKB approach, at least at k • B 0 = 0. Therefore, one should treat the results of papers which do not show the splitting with care. In section 2.1 we use the results of [1.61.11, 2.52.7]. The shortwavelength perturbations with kz — 0 discussed in sec16
Simplified Dispersion Relations
17
tion 2 . 2 can be of importance in the study of lowfrequency instabili ties where a; < fli, for ioncyclotron instabilities where u ~ £}*, and for highfrequency ones where u Ct[. The main results of section 2 . 2 were obtained in [2.8, 2.9]. The problem of shortwavelength perturbations with u>  kzU)/ \ kz  vt , & = u  kzU. The relations in (2.26) were obtained in [2.7]. They will be used in Chapter 13. 2.1.6
Allowance for the interaction of Alfven perturbations with resonant particles In contrast to (2.8), equation (2.7) does not contain the terms cor responding to resonant interaction between waves and particles. In order to take into account such an interaction we should, first, aug ment cn by terms of order f 2 and, second, allow for the contribution in (2.4) by A. Then (2.7) is replaced by Di0) + 4V + A (0 )/D 2 = 0
(2.27)
where fl11') is a correction to tn of order £2, and A^°* denotes the leading terms of expansion of A in powers of k \ . For subsequent calculations of the growth rate due to resonant in teraction, we need to know only the imaginary part of According to (2 . 1 ), it is equal to
i me “
( 228)
Note: the righthand side of this equation represents the contribution due to the ions; the ion indices are here omitted for simplicity; as before, S± = v \ j 2 . We find the expression for A^°\ taking into account equation (2.7), in the zeroth approximation in £ 2 A =  C XD 2  C2.
(2.29)
Simplified Dispersion Relations
25
Here (2.30)
(2.31) Then equation (2.27) reduces to D (°} +
 Ci  C2/D 2 = 0.
(2.32)
Using (2.32), we find the general expression for the growth rate 7 = Imu; of the resonant interaction between the Alfven perturbations and the particles ^ Im (€ $  C 1  C2/D 2).
(2.33)
When kz = 0 , equation (2.33) is replaced by (2.34) Relations (2.33) and (2.34) will be used in Chapters 4 and 5.
2.2
D ispersion relations for shortwavelength perturbations with kz — 0
Let us now consider how the general dispersion relations will be sim plified if the transverse wavelength of perturbations is small com pared with the ion Larmor radius, k±p\ > 1, but larger than or comparable with the electron one, k±pe < 1 . The perturbation fre quencies are assumed to be small compared with the electron cy clotron frequency, u l. (3.51) Note also that, for such 77 and j3 is important. This means that the ion longitudinal thermal motion is assumed to be significant. Generalization of equation (3.28) to the case of arbitrary kzvTJw is the following equation obtained by using relations (2 .8 ), (2 .6 ), and (2.9):
+T4 m ^&i{I 1,(\\ nm"M w kz \vTJ li ( VIrirr~ h I vTi {x2w) hi
^H D =o
UDi kz
(3.52)
Here W (x) is defined by (2.24) where x = (u> —u>DiA)/ kz vTi
(3.53)
and A = Miv]_/2T. The symbol (...) means averaging only over the transverse velocities, this symbol is defined in equation (2.19). For simplicity, in obtaining (3.52), we restricted ourselves to the case Te = T, = T. In the low/? limit, when the terms with fi and wD; are unimpor tant, equation (3.52) reduces to
Electromagnetic Instabilities in an Inhomogeneous Plasma
48
This equation describes the wellknown ion temperaturegradient drift instability (see [2], section 3.2) originally studied in [3.223.24]. In the discussed longwavelength approximation, such an instability takes place for r) < 0 and 77 > 2. (In the case of perturbations with k±pi ~ 1 , the boundary of this instability in a plasma with positive 77 is displaced to 77 = 0.96, [3.24], see also [1.3].) The results following from (3.54) for u > kzvTi9  77 —► 0 0 and for u 2. The present analysis has two goals. First, we shall study how equations (3.55) and (3.56) are modified when accounting for the effects of order (3 in the approximation (3 0, the phase velocity of the perturbations with = u j + is directed towards the electron magnetic drift, uj+/uj^e > 0 , while that of the perturbations with uj = uj is towards the ion magnetic drift, uj^/ujBi > 0 . Under the condition
uj
0(1  2/f3)2/8 < a < 0
(4.7)
[1 + 20/(1 + 0/2 ) 2] ' 1 < 1 < 1/2
(4.8)
and (3 < 2 , both waves propagate towards the electron magnetic drift, u;±/u;De > 0. When (3 > 2, they propagate towards the ion magnetic drift, u j ± / u jD i > 0. Waves with u; = for (3 < 2 as well as waves with u = for (3 > 2 have a negative energy u d e ^ /d u < 0.
(4.9)
The above results will be used in the analysis of the resonant interaction between Alfven waves and particles. 412 Perturbations with kz ^ 0 For kz / 0, it follows from (2.5), (2.7), and (2.9) that (4.1) is replaced by u 2+ UJDi) +  k2 zc\= 0. (4.10) Neglecting the gradient terms, we find from (4.10) the dispersion relation for the Alfven waves in a homogeneous plasma a, 2 = k]ci .
(4.11)
It follows from (4.10) that, for kz ^ 0, the instability of a plasma with rj lying in the range of (4.2) is suppressed. Qualitatively, the stability condition means < kzcA.
(4.12)
At the same time, condition (4.12) characterizes the order of magni tude of maximum kz when the gradient effects are still important.
Alfven Magnetodrift Instabilities in a FinitePlasm a
4.2
57
Interactions of driftAlfven waves w ith resonant particles for kz = 0
42.1 Resonant interaction of the waves with electrons Using (2.34), we find that interaction between Alfven waves with kz — 0 and electrons is characterized by the growth rate 1 £12 7= — d e ^/d u I e22  N* P
•(e)
22
(4.13)
where, according to (2.3),
4 WDe
P 1 + 7?
u
  e/ )• (4.14)
(aj Waves with u = a>+ According to section 4.1.1, electrons interact with waves of type u — u>+ for all P if a > 0, and for P < 2 if a satisfies condition (4.7). Using (4.13) and (4.14), we then obtain the instability condition 2 7? /? 1 + 77
417? p i + 7?
This condition is not satisfied for any 7? ^ 0, i.e., the perturbations discussed cannot be excited in plasma with 7? > 0. However, an instability is possible for 7? < 0. In particular, when P C l , condition (4.15) means (1  2/3) < 7? <  p . (4.16) This range narrows with increasing P, and the instability disappears when P = 2. (b) Waves with u = u>As stated in section 4.1.1, electrons interact also with waves of type u = u;_ under condition (4.7) and when P < 2. Allowing for (4.4), we find from (4.13) and (4.14) the following instability condition for these waves 4 1— + 1+1/n +i'/n +J ' K r 1) +j ] +M 7?
>0'(4'17>
The equation for the instability boundary for low P following from (4.17) is of the form 3t?2 + rj —1 = 0. (418)
58
Electromagnetic Instabilities in an Inhomogeneous Plasma
The condition a < 0 for low (3 means —1 < rj < —0.5. Hence we find that a resonant instability occurs for low j3 if (4.19)
—0.77 < rj < 0.5.
As (3 increases, the range of the parameters 77, corresponding to the instability, narrows. When f3 = 2, this range disappears, as well as that discussed in section 4.2.1.(a). 42.2 Resonant interaction of waves with ions We start with the general relation for the growth rate of the Alfven perturbations with kz = 0 following from (2.34): (4.20) Here a i 2 = —i^i25 the superscript (1) in Imen is omitted for sim plicity. It follows from (2.3) and (2.28) that Im eu = X/4 Im a ^ = ftjX/a;
(4.21)
Ime22 = 4(1?X / u 2 where
_ fa;nj(l ~ ??) 2 u
1}
(4.22)
P l +V
Allowing for (4.21) and (4.22), we reduce (4.20) to the form X O ')/*
d $ /d u
uj
[(l/2)(Ree22  N 2)  2 (n ju )(K e a 12)}2 I e22 
I ,
NM 22 122
•
V1  ^ )
Obviously, using this equation, we should allow for the frequencies u j ± to be real, i.e., rj should satisfy inequalities (4.6) or (4.8).
=
(a) Waves with uj — u j _ In the case of waves where uj = a;_, equations (4.22) and (4.23) yield the following instability condition (cf. equation (4.17))
59
Alfven Magnetodrift Instabilities in a Finite/? Plasma
In particular, it follows from (4.24) that there is no instability for low f3 and permissible values of rj. In addition, assuming (3 ^ 1 , we find that in this case there is also no instability for permissible rj. Therefore, we conclude that, for all values of the parameters (3 and 77, the interaction between the wave of type and ions does not lead to instability. (b) Waves with u = u;+ The ions interact with these waves for (3 > 2 and when a satisfies condition (4.7). Allowing for (4.4), (4.22) and (4.23), we find that in this case the instability condition has the form
When /? >
1
, this condition together with (4.8) means (18//3) < r)< 1/2.
(4.26)
Thus, the interaction between waves of type w+ and the ions in a plasma with /3 > 2 results in an instability in the range of negative values of the parameter rj determined (for 3 3 > 1) by condition (4.26). As d decreases to /? = 2, this range narrows and, when 3 = 2 , the instability disappears. 4.3 43.1
Role of resonant particles in Alfven perturbations w ith kzvTi < u < kzvTe
The zeroth approximation in the parameters u /kzvTe and kzvTi/w The Alfven perturbations with kzv?, kzvre in a plasma with f3 ~ 1 in the zeroth approximation for parameters z\and (kzVTi/u>)2, as well as for kz = 0, are described by the dispersion relation (4.1). The frequencies of these perturbations were given in section 4.1.1. For finite kz, resonant interaction between the waves and the parti cles is modified in comparison with section 4.2. Under the condition u> /kzvTe this interaction can be ignored; this case is dealt with in this section. However, such an interaction will be taken into account in section 4.3.2. The condition of resonant interaction between waves with finite kz and the ions means ui = kzvz + o>di. When the ratio kzvTi/u>
60
Electromagnetic Instabilities in an Inhomogeneous Plasma
is small, the term with kz in this condition is essential only for an exponentially small number of particles. In this section we neglect such an effect and consider it later in section 4.3.3. Under these assumptions, the general problem of resonant inter action between the Alfven perturbations and the particles reduces to the analysis of the resonance of type u = The growth rate of the Alfven perturbations due to this resonance is given by the expression following from (2.33):
7
=
(e,‘ ^ / /J 4
(4'27)
where a 23 =  ^ 2 3 • Since the value a u = —iei2 depends solely on ions (see (2.3)), its form turns out to be the same as in the case where kz — 0. On the other hand, the values of a 23 and 633 under the above assumption (kzVTi < w < kzVTe) are determined by the electrons. These quan tities are real when the terms of order u/kzVTe are neglected. As for component e22, the contribution of the electrons to it in the zeroth approximation in u /kzvTe is vanishingly small, and that of the ions is the same as in the case where kz = 0. Then, using relations (4.21), we reduce (4.27) to the form similar to (4.23) X [(l/2)(Re e22  c^3/e33  N 2)  2(Oi/a;)Re « 12]2 de^/dcj I £22 — £*23/633  N 2 2 (4.28) Hence it can be seen that the problem of determining the sign of the growth rate of resonant interaction reduces to that discussed in sec tion 4.2.2. Therefore, all the conclusions of section 4.2.2 concerning the resonant instability conditions holds in the case discussed. It is also evident that the absolute value of the growth rate for kz > u/vTe is different from that in the case where kz = 0. However, we shall not dwell upon the question of the absolute value of the growth rate since, for the most interesting values of the parameters r/ and /?, it is negative. 43.2 Interaction of waves with resonant electrons Using (2.3), we evaluate the quantities D 2,C i and C2 determined by equations (2.6), (2.30) and (2.31) and conclude that, first, the contribution of C\ into the general expression for the growth rate (2.33) can be neglected, and, second, the terms 6 ^ and 613 in the expression for C2 can be omitted. A small parameter which allows
Alfven Magnetodrift Instabilities in a Finite/} Plasma.
61
one such an approximation is (kzVTi/u)2. Then, we take into account the fact that electron resonance is important only for perturbations propagating against the ion magnetic drift, uj/ ujy>\< 0 , since other wise the electrons would produce only a small correction to the ion growth rate. For such perturbations the quantity a i 2 = —i^i2 &s well as the ion contribution to 622 a*e real. The resultant expression for the electron growth rate is /du\ ^2N2  a l 3/e33\ 2
7
(4 29') >
As in equation (4.27), the quantities a 23 and 633 used here include only the electron contribution. But now, in contrast to (4.27), imag inary terms are also allowed for in these quantities. In addition, the imaginary contribution of electrons to e22 should be taken into account in deriving (4.29). In this adopted approximation, the quantities ,a 23, and e33 are determined by equations (2.62)(2.65). Using these equations, we find lm
_ !k\ = c 33 /
V
2
I kz I vTe g2
(4.30)
where (. =
W„ e \ 3
3 Wne/.
I1—J
W„e\ (4.31)
 v — {1  — J + j ^ UJne
U>De
UJ
UJ
g2 = 1  ( l  ^ ( l + , ) ) .
(4.32)
The instability condition following from equations (4.29) and (4.30) has the form gtdefl/du < 0.
(4.33)
This condition is not satisfied if 77 ^ 0, even for large positive rj. However, there is an instability for large negative rj; in this case the plasma turns out to be unstable for both low and high /3. 4.3.3 Ion resonance of the type u> = kzvz + wdj for w/wdi < 0 In order to find out how the general expression for the growth rate (2.34) is simplified in the case of ion resonance of type u = kzvz +u>di
62
Electromagnetic Instabilities in an Inhomogeneous Plasma
for cj/ujdi < 0 , we shall consider the following integrals over velocities in the antiHermitian parts of eap I ns =
J
F\n(vz/vT)s6(u)  kzvz  (jjDX)d£j_dvz
(4.34)
where A = ME^/T^n = 0 , 1 , 2 ,3,4;s = 0 , 1 , 2 ; and F is the Maxwellian function of (1.31). The indices are omitted for simplicity. After integrating over vz, we obtain nll2no In‘
(
u2 \
( u>  u>D\\
VTfc,eXPl X exp
kzvT
_AI 1  2u;p^ "\ _ A2u;d kyTJ
l/rp vt vi
J
dA.
In this calculation we assumed u and uv to be values of the same order, but opposite in sign, < 0 , and that ( , u; ) > kzvj>. Consequently, a substantial contribution to the integrals of (4.35) is due only to sufficiently small values of A A < k*v%/ 
< 1 .
(4.36)
Therefore, the integrals I ns reduce to 7r1/2n 0 (
____ (
u
Vt x
u 2
^
eXPV k2 zv\)
f°° \n J.
* “H
(
2A I WdW
(4.37)
" 
It is also clear that the integrals with the subscripts n > 0 contain additional small factors of order (k2v u;Du; )n whilst the greatest among the integrals with subscripts s = 0 , 1 , 2 is the integral with s = 2. Therefore, it is sufficient to take into account the contribution to the growth rate due only to the component eap containing I ns with the smallest value of n and the greatest value of 5 , i.e., to the component £33(71 = 0 ,s = 2 ). Then, using equations (2 .3 ),(2.34) and (4.37), we obtain
7=
 
de^/duj [(&>  iV2)Re e33  (Re a 23)2]2
(4 .38)
where T (i) X ul\exp(—u; /A;2^ ) Im ?3’ =  T T rrSa I kz vTiuDi
1 W„i f
3\
U>U>Ti
* = 1 i r l 1 P )  &>;■
(4.39) , .
...
(4A0>
63
Alfven Magnetodrift Instabilities in a Finite3 Plasma
The quantities Re a 23 and Re e33 are determined by (2.62) and (2.63), while the real quantities a 12 and 622 are calculated by means of (2 .3 ) for kz = 0 . From (4.38) we obtain the instability condition  ^ ig s d tn /d u < 0.
(4.41)
In particular, it follows hence that the instability is absent when 77 = 0 , but that it does occur at arbitrary small positive 77 77
> 0.
(4.42)
When the assumption u /k2vTl > 1 is used, the growth rate is exponentially small, but it becomes significant at the validity limits of this assumption. Note that, when 77 > 0, perturbations with < 0 correspond to the branch with u = u;+ determined by equation (4.4).
4.4
Role of resonant particles in Alfven perturbations w ith u ~ kzvT\for VT = 0
Since expression (2.33) for the growth rate of Alfven perturba tions with u ~ kzVTi is very complicated, we restrict ourselves to analysing the resonant excitation of such perturbations in a plasma with VT = 0. Even with this simplifying assumption, the problem still remains too complicated. Therefore, we also assume that the electron temperature is lower than the ion temperature, Te d) + Im (^(3 + u,dCo)^’d) \ aQqKzVz) and Reu; is determined by (5.31), i.e., in the explicit form Rea; = ±kzcA(l + 2u^/k2v^)1/2.
(5.36)
Since A does not depend on u (see (5.32)), it can be seen from (5.34) that one of the oscillation branches is unstable, 7 > 0. An ad ditional analysis similar to that in section 4.4 shows that the unstable branch is that with Keuj/ujn > 0. This fact is natural since otherwise the imaginary parts of eap have the same signs as in the case of a homogeneous plasma, while the energy of the waves considered is positive. In order of magnitude 7
~ (3(kxp fc A/a
(5.37)
where p is the characteristic Larmor radius of highenergy particles and a is the characteristic length of their inhomogeneity. Note that the analysis of the same problem in [1.10, 5.5] is not precise. In the mentioned papers only the term Im en was taken into account, while the term Im[ef2 /(e22 — N 2)] in (5.3) was neglected assuming it to be small, of order k\p2, in comparison with Im en. In reality, both terms have the same order of magnitude. It is clear that this inaccuracy is quantitative but not qualitative. Therefore, the qualitative results of [1 .1 0 , 5 .5 ] remain valid and are in accordance with the ones reached above. 5.3.2 Instability of the shortwavelength Alfven perturbations According to (5.37), the growth rate of Alfven waves increases with increasing k±p. However, for k±p ~ 1 expression (5.34) and the estimate (5.37) become invalid. In order to study the excitation of Alfven waves for k±p > 1 we shall use the general dispersion relation (1.1) with eap of the form of (2 .1 ) and (2.2). As has been considered throughout this chapter, we assume e33 to be sufficiently large, so that equation (1.1) reduces to (5.1) again. Using the asymptotics of the Bessel functions for large arguments, we then find that the
78
Electromagnetic Instabilities in an Inhomogeneous Plasma
components e12 and e2i are unimportant. As a result, we arrive at the following dispersion relation similar to (2.7): Cn  k2c2/u 2 = 0.
(5.38)
We then obtain the same expression for the real part of the fre quency as in the case of a homogeneous plasma Rew = ± kz cA.
(5.39)
The growth rate is equal to 7
T /(d F
, ky 8F\ u>d 2A c,
,
/c ^
Here the notations are the same as those adopted in (2.2). The result of (5.40) holds for an arbitrary ratio vT/cA. If, by analogy with section 5.3.1, we adopt vT > cA, it follows from (5.40) that
7=
+
“d)sg“(«) )•
(541)
In order of magnitude 7
~ flcA/a(k±p)2
(5.42)
so that the growth rate decreases as kJ 2 with increasing k±. It is clear from (5.41) that the branch with Reo;/o;n > 0 is unstable (see section 5.3.1). 5.3.3 Summary and discussion of the results At the validity limits of the qualitative relations (5.37) and (5.42), i.e., at k±p ^ 1 , these relations reduce to the same estimate: 7  7max ^ f}cA/a.
(5.43)
The longitudinal wave number of such perturbations is of order kz  vTf3/cAa.
(5.44)
It should be kept in mind that for k±p ~ 1 the wave frequency becomes of the order of the growth rate Reu; ~
7
~ ficA/a.
(5.45)
Lowfrequency Instabilities in a Twoenergy Component Plasma
79
Hence the method of expanding in a series in the small parameter /Re a;, used in the above, proves to be at the validity limits for such k±p. Thus, it follows from our analysis that in a plasma with a group of highenergy ions with an inhomogeneous density, the Alfven waves can be excited. The characteristic transverse wavelength is of the order of the Larmor radius of these ions. The growth rate and the frequency of the unstable perturbations have the same order of mag nitude determined by (5.45). The longitudinal wavelength is given by (5.44). The Alfven growth rate (5.45) is small compared with the magne toacoustic growth rate (5.12). However, the Alfven instability does not require the threshold of (5.7). Therefore, under condition (5.7) one can restrict oneself only to the allowance for the magnetoacous tic instability, while under the opposite condition to (5.7) only the Alfven instability is possible. 7
6
Electromagnetic Instabilities in a Strongly Inhomogeneous Plasma
For a finite ratio of the Larmor radius of particles to the charac teristic length of a plasma inhomogeneity, it is necessary to allow for the cyclotron and highfrequency instabilities which are the sub ject of this chapter. Under the condition where the ion tempera ture is large or at least not too small compared with the electron temperature, such instabilities are associated with shortwavelength perturbations (with respect to the ion Larmor radius). In this con nection, we start by analysing the shortwavelength perturbations of a Maxwellian plasma neglecting the cyclotron and highfrequency resonances and assuming kz = 0, and E ± B q. This is the goal of section 6.1. Then we shall examine finite/? effects in the driftcyclotron instability (section 6.2), the lowerhybriddrift instability (section 6.3), the drift losscone instability (section 6.4), and the ion cyclotron excitation of the transverse drift waves (section 6.5). As in section 6.1, we assume kz = 0 in sections 6.26.5. The study of instabilities with uj/kz < vTe is also of interest. This is the goal of sections 6.6 and 6.7 where the instabilities associated with shortwavelength driftAlfven oscillation branches axe consid ered. It is assumed in section 6.6 that u (MejM i)1^2, and (3 ~ 1, the instability is suppressed for the highest harmonics but holds for harmonics with small numbers. However, in the latter case the growth rate of the instability is a few times smaller than that when /? 1. An examination of the role of increasing the electron temperature, i.e. /3e, was made in [2.9]. The results of this paper are presented in section 6.4.2. It emerges that when the electrons are heated to the extent that (3e ~ 1, the drift losscone instability caused by ions with f3i > 1 may be totally suppressed. Evidence of the drift losscone instability in experiments on the confinement of hotion plasma was reported in [6.306.36] and others. One of the main and experimentally verified methods of suppres sion of the drift losscone instability is the injection of a plasma of a relatively low density with slightly warm ions [6.306.33]. Such a method is called “stabilization by warm plasma”. In [1.1] and then in [6.26] the problem of the stabilizing effect of increasing (3 in com bination with the injection of warm plasma on this instability was analysed. This problem is discussed in section 6.3.3. In particular, it turns out that the larger the value of faxthe smaller the warmplasma temperature necessary for stabilization (see inequality (6.64)). A possibility of the existence of the drift losscone instability in the tokamak divertor was considered in [6.37]. The role of the drift losscone instability in magnetospheric plasma was discussed in [6.29]. According to [6.29], certain varieties of this instability can play an important role in the dynamics of the ring current and the inner edge of the plasma sheet region of the mag
84
Electromagnetic Instabilities in an Inhomogeneous Plasma
netosphere, and would give rise to precipitation of protons on the auroral field lines, which may contribute to the excitation of the dif fuse aurora. In addition, the authors of [6.29] consider that these instabilities may be relevant to the observation of the ioncyclotron harmonics by [6.38]. In contrast to the instabilities presented in sections 6.26.4 relat ing to hotion plasma, the instabilities studied in section 6.5 can be excited in a strongly inhomogeneous hotelectron plasma. The the ory of these instabilities was developed in [2.5] and further in [3.8, 3.9, 3.21]. Space applications of this theory were given in [3.193.21]. As mentioned in Chapter 3, these applications concern the problem of electromagnetic noise in the hydromagnetic discontinuities of the solar wind. The shortwavelength driftAlfven ( s d a ) waves discussed in sec tions 6 . 6 and 6.7 were initially studied in connection with the prob lem of anomalous plasma transport [2.10, 6.39]. It was then assumed that their frequencies were small compared with the ion cyclotron frequency. Subsequently the authors of [6.406.42] have noted the importance of s d a waves in the problem of highfrequency heating of plasma and extended the notions about these waves into the re gion of frequencies higher than the ion cyclotron frequency, u >> 11*. According to [6.406.42], s d a waves with u > fi* can be excited by a transverse current arising from the influence on plasma of an alternating electromagnetic field (see in detail Chapters 7 and 8 ). As a result, such an instability can lead to fast heating by this cur rent. Recently the notions about s d a waves were used to explain anomalous electron heat conductivity in tokamaks [6.43, 6.44]. This is associated with the fact that s d a waves result in bending of the magnetic field lines, so that the electrons moving along such field lines can be displaced across the equilibrium magnetic surfaces. Fur ther development of the nonlinear theory of s d a waves can be found in [2.11, 6.456.47].
6.1
Shortwavelength perturbations in a finite/? Maxwellian plasma
Using the results of section 2.2.1, we shall consider shortwavelength (k±pi > 1) perturbations with kz = 0 and E ±B 0 in a finitef3 Maxwellian plasma. In section 6.1.1 the perturbations are assumed to be of long wave length, with respect to the electron Larmor radius (k±pe —>0), while in section 6 .1 . 2 the finiteness of k±pe is taken into account. In both
Electromagnetic Instabilities—Strongly Inhomogeneous Plasma
85
cases we shall restrict ourselves to the analysis of a plasma with Te = T[ = T. However, in section 6.1.3 we shall assume Te = 0. 6.1.1 The k±pe 0 approximation Under the abovementioned assumptions and with k±pe —> 0 , the perturbations are described by dispersion relation (2.35) with 6 n = and £12 = where and are defined by equations (2.42), while 622 is of the form of (2.43). Substituting the abovementioned expressions for eap into (2.35), we arrive at the dispersion relation
Note that within the accuracy of an interchange of the species subscripts, equation (6.1) coincides with equation (3.28) taken for Te = Ti. Therefore, to find physical results from (6.1) we can use the analysis given in section 3.3. Then we conclude that equation (6.1) describes two oscillation branches: the branch of iondrift waves (iondrift oscillation branch) and the magnetodrift branch. In the absence of a temperature gradient, VT = 0, the frequency of the iondrift oscillation branch for arbitrary /? is uj =
u nl
( 6 .2 )
so that in this case (i.e., for VT = 0) the oscillation frequency is real. For small VT( 77 < 1 ) the solution of equation (6.1) for this branch is (cf. (3.39)) ( i , _ Z3/2 W ” a’"i V i + /? /2
rf P \ 4 (l + /3/2)3 (CoeA)J
(a o\ ( • }
where A = Mev\/2T, (0e = (u  u^e)”1. It can be seen that un der these conditions the oscillations are damped with a decay rate proportional to rj2 due to their interaction with resonant electrons. The shortwavelength magnetodrift oscillation branch for 0 < ?72 — 1 ^ 1 is characterized by the frequency (cf. (3.42)) 1+V UJ = —U De e x p
(6.4)
Resonant interaction of such perturbations with electrons is absent, while that with ions is negligibly small due to the adopted approx imation k±pi  * 00. In a plasma with /? < 1 and 1 De/2 .
(6.5)
86
Electromagnetic Instabilities in an Inhomogeneous Plasma
For 7] < 0 these waves can interact with resonant electrons, but, since  uj > UDe, the growth rate of such an interaction is exponentially sm all
6.1.2 Electron damping of the iondrift oscillation branch According to the above, when rj = 0 and k±pe —» 0, in a plasma there is an iondrift oscillation branch (6 .2 ) with zero growth rate. The phase velocity of such oscillations, u/ky, has the same direction as the electron magnetic drift. Therefore, if small terms proportional to high powers of k±pe are taken into account, as well as for r] ^ 0 we must obtain an imaginary addition to the frequency due to the interaction of the waves with resonant electrons. Since the operator le for rj = 0 is equal to 1 —cJne/uJ, the sign of le in the case of u j « uon \= u jne is positive. Hence as in the case of a homogeneous plasma, resonant electrons remove energy from the waves, ujlme^2 > 0, and the waves must be damped, 7 < 0. Let us verify this by a straightforward calculation of the growth rate (i.e., in this case, of the decay rate). We take into account that, in the zeroth approximation in k±pe, the solution of equation (2.35) has the form of (6 .2 ). Then, according to (2.42), = 0 (we allow for VT = 0). Therefore, to an accuracy of the terms of order (k±pe ) 2 inclusively, equation (2.35) reduces to the form eft + eft = 0 (6 .6 ) —
where eft is defined by the first equation of (2.42), and eft is defined by the first equation of (2.41). It follows from the mentioned relationships that w = « ni[l  2ze(l  0 /2)(1 + 0 /2)"1].
(6.7)
It can be seen that the addition to the oscillation frequency of order 2 e is real. Hence, to find the growth rate (or decay rate) of the oscillations it is necessary to take into account the terms of order (kxPef. Starting with equation (3.6) and allowing for the abovementioned smallness of eft for VT = 0, we find a general expression for the growth rate (cf. (2.34) and (4.20)):
Electromagnetic Instabilities—Strongly Inhomogeneous Plasma
87
We use the fact that (cf. (4.21)) I m e f f = X /4
Ini (—if 12) = Q.eX /u
(69)
Ime 22 = 4X (fte/u>)2 where in the case discussed 1T Z2
U3
(
U)
With the help of (6.9) we transform (6 .8 ) into 1
a
7 = “ S S ? /fc
< 6 ' 1 1 )
where a 12 = i(c(a2) + eli l )  (c22  N 2)zeu>KB/4uDek = (fcKBA^e)_1ze/?(l  p/2).
(6 .12)
Since code^/du = 2, u;Ime22 > 0, it can be seen from (6 .1 1 ) that, in accordance with the above, 7 < 0 , i.e., the oscillations are damped. When fi ~ 1 , it follows from (6.11) that the decay rate is of order 7
~  Rea; z\.
(6.13)
In the case where (3 1 , the approximate formula J 2 (£i) ~ (^fi ) 1 *s true (see also section 2.2.1). Using this formula, we calculate the average in (6.31) and arrive at the following expression for the growth rate ~=
U>2 ( 1  u>ni/u>) exp [—(a;  nnQ/cjpj] I wDi [2  Io(ze) exp(—ze)] [2zi(u  n t y ) / ^ ] 1/2'
,
.
The growth rate should satisfy the above condition 7 < u — nQj. This yields a lower limit for the difference u —nfi\entering in (6.32). Hence an estimation of the maximum growth rate can be obtained by assuming 7 ~ uj — n il[. In this case we have w — nfii Di, so that the exponential in (6.32) may be replaced by unity. We then obtain
When p j a ~ (M e/M i)1^2, only perturbations with u> ~ J)j(n = 1) can be excited. In this case it follows from (6.33) that Tmax *
(6.34)
At the validity limits of inequality (6.29), i.e., when f3 ~ (Me/M;)1/4, equation (6.34) is replaced by equation (6.22). When p j a > (M e/M i)^ 2, it follows from (6.33) that for the high
est harmonics
7max *
(6.35)
while for the first harmonic 7max ^
Sl\(Mea/P l/2MiPif13.
(6.36)
Similarly to (6 .3 4 ), when (3 ~ (3min, expressions (6.35) and (6.36) are transformed to the respective relationships of the approximation (3 0 (see [2], section 3.1). The instability considered here can be called the electrostatic magnetodrift driftcyclotron instability. 6.2.2 Electromagnetic driftcyclotron instability W ith increasing /?, the electromagnetic part of the perturbations and the electron magnetic drift become important. These effects are taken into account by equation (2.35) with the components eap given in section 2.2.1. Now we proceed to the analysis of this equation. The corresponding instability can be called the electromagnetic driftcyclotron instability.
92
Electromagnetic Instabilities in an Inhomogeneous Plasma
In section 6.1 we analysed equation (2.35) neglecting the ion res onance. It was shown that the perturbations were damped due to their interaction with the resonant electrons. The decay rate was expressed by the approximate relationships (6.13)(6.15). Besides the electron damping, there is one more effect associated with the electrons and revealed at finite /?; this effect is described by equation (6.7) and is the change of sign of the value l —ujni/u>. This sign becomes positive when (3 >
2
.
(6.37)
Then sgn(u;Im en ) > 0, so that the ions do not result in an excitation, but in a (lamping of the perturbations. Hereinafter we shall assume (3 < 2 and examine under which con ditions electron damping overcomes ion excitation. (a) Stabilization of perturbations for p ja > (M e/M i ) ly^2 In this case ze — 1 , so that the electron decay rate is of order (6.14) while the ion growth rate is of order (6.34). Comparing these two expressions, we conclude that stabilization takes place even for small P (3 > 3[ln(Mi/M e ) ]  1 « 0.3. (6.38) (b) Stabilization of the highest harmonics at p ja (Me/Mi) 1/ 2 As in section 6 .2 .2 .(a), here we also have — 1? and (6.14) still holds, while the ion growth rate is now determined by (6.35). As a result, criterion (6.38) still holds. (c) Perturbations with uj ~ ft* for p j a > (M e/M \ )^ 2 Now *e ~ (a/Pi)2M e/Mi < 1
(6.39)
so that we have estimates (6.13) and (6.15) for the electron decay rate. For (3 < 1 the ion growth rate is of the order of (6.36). When P ~ 1 , it is determined by the dispersion relation of the form of (6.19) where e ^ , 4V, an(l &€n are given by (2.41) and (2.42). Solving (6.19), we find that (6.36) holds up to (3 ~ 1 (but, in reality, it is considerably modified if (3 > 2, see condition (6.37)). Due to the smallness of the electron decay rate (6.15), for small j3 stabilization does not take place, so that the instability occurs up to (3 « 2 .
Electromagnetic Instabilities—Strongly Inhomogeneous Plasma
6.3
93
Electromagnetic lowerhybriddrift (highfrequency drift) instability in a finite/? plasma
As mentioned in section 6.2, under condition (6.27), the driftcyclotron instability in a zero/? plasma turns into the lowerhybriddrift instability. Then the growth rate of separate harmonic becomes of the order of the cyclotron frequency and, as a result, a substantial broadening of the cyclotron resonance occurs. The broadening of the cyclotron resonance is also caused by ion magnetic drift. For a sufficiently large ion magnetic drift, i.e., for sufficiently high /?, the frequency range of such a broadening can also be compared with the ion cyclotron frequency, i.e., u>di >
(6.40)
As mentioned in section 2.2.1, in this case it is necessary to use the approximation of a continuous spectrum of cyclotron harmonics, i.e., the socalled highfrequency approximation. Let us consider under what /? this approximation is valid. We allow for magnetodrift broadening of the cyclotron resonance to be due to particles with thermal velocities, i.e., with v]_ ~ 7]/Mi, in contrast to the situation considered in section 6 .2 . 1 when v\ ( M j M ^ a / p i .
(6.41)
Comparing this inequality with (6.28) and keeping in mind (6.27), we conclude that the condition /? > (Me/Mi)1' 4, treated in section 6 . 2 as a condition of importance for the ion magnetic drift in the driftcyclotron instability, can also be treated as a condition that the broadening due to the ion magnetic drift prevails upon the broaden ing due to the growth rate. In other words, under condition (6.29), the transition to the highfrequency approximation is determined not by condition (6.27), which can be invalid, but by condition (6.41). If /? becomes of order unity, it can be seen from (6.41) that condi tion (6.27) is replaced by a condition of type (6.21). In other words, when /3 ~ 1 , the instability of a plasma with finite ion Larmor radius caused by density inhomogeneity resembles a highfrequency insta bility, but not a cyclotron instability. Of course, some features of a cyclotron type of instability should be revealed. Furthermore, in the case of a strongly inhomogeneous plasma our estimates correspond to the highest frequency range of the spectrum associated with rather shortwavelength perturbations. A longer wavelength and lower fre quency range of the spectrum must remain substantially cyclotronic even though condition (6.41) will be valid.
94
Electromagnetic Instabilities in an Inhomogeneous Plasma
Based upon the above, we shall assume condition (6.41) to be valid and examine the lowerhybriddrift instability at first in the approximation of cold electrons, Te = 0 (section 6.3.1), and then by allowing for finite Te (section 6.3.2). 6.3.1 Coldelectron plasma For Te = 0 (coldelectron plasma) the problem of lowerhybriddrift instability is reduced to the allowance for ion resonant interaction described by 2(M e/Mi)upe/c.
(6.52)
This is the boundary of the drift losscone instability in a plasma with fi > (Me/Mi)1/3. Condition (6.52) can be written in a form similar to (6.46) p j a > (M e/ M ) 1/2/?1/2.
(653)
It can be seen that the critical value of the relative density gradient 1 /a, for the onset of plasma instability, increases with /?, going as fl1/2. This situation can be regarded as being favorable for stable plasma confinement. When (3 ~ 1, the density gradient must be such that (cf. (6 .2 1 )) Pi/a > (Me/Mi)1/2.
(6.54)
The growth rate is reduced due to the electromagnetic nature of the perturbations. When fi > (M e/M [Y^ and p ja is not very much larger than the righthand side of (6.53), we find
( 6
m
)
In contrast to the case where fi + 0, the growth rate is small compared with the ioncyclotron frequency. 6.42 Hotelectron plasma Now we consider fiQto be finite (hotelectron plasma). For simplicity we assume fii > 1 . We describe the resonant interaction of ions with waves by means of (2.56). Let us study separately the cases where fie < 1 and fie > 1 , using the approximate relationships from section 2 .2 .1 . The electron temperature gradient will be neglected. When fie < 1 and fii > 1 , the real and imaginary parts of the wave frequency become the following (cf. (6.18) and (6.42)): Recj = o>nj(l j2/fi) 4U2 (
kv0fii \ where
7T
(6.56)
1 \ 7 = T r i H 2 \kb \pi J1  o i (65 7 )
97
Electromagnetic Instabilities—Strongly Inhomogeneous Plasma
(6.58)
$ (A 0 = (2 //?e) 2 exp(—2//3e).
The first term in the large brackets of the righthand side of (6.57) describes the ion excitation, while the second describes the electron damping. Recall that our analysis concerns the case of a weakly inhomoge neous magnetic field, kb px 1 and, as above, $ > 1. In this case we have (6.59) )
(6.60)
where 7] = M[vl/2. It can be seen that, when Te < Ti, the pertur bations are damped for all values of kb Pi When Te > T*, it is a necessary condition for stabilization that I kb I Pi < T\/Te.
(6.61)
Thus, with the objective of the suppression of drift losscone in stability, too high a heating of electrons (Te > Ti) for /? > 1 is undesirable. 6.4.3 Stabilization by a warm plasma In the presence of a small admixture of slightly warm Maxwellian ions, equation (6.48) is replaced by Do + SDq — 0
(6.62)
where D 0 is the lefthand side of (6.48), while 6D0 is of the form
n
a;J„(zi)exp(zi) u; —nQ,j
(6.63)
Here a = 7ii/ n 0 < 1, is the number density of slightly warm ions, p\ = Ti/M\ili is their squared Larmor radius, Tx is their tempera ture, and z\= (kpi)2. Using (6.62) and (6.63), we can evaluate the effect of the warm ions on the drift losscone instability in the case where the density gradient is not very large, so that inequality (6.53) is not strong. In
98
Electromagnetic Instabilities in an Inhomogeneous Plasma
this case the characteristic value of k is of the order of cjpe/c so that zi ~ /?Ti/Met>o . Let Tx > M ev2 0/I3. (6.64) In this case z\> 1 so that I n(zi) exp(—zx) « (2irzi)1^2. Using these estimates, we find by means of (6.62) and (6.63) that the instability is impossible if a > t (6. 65) When j3 ~ 0.1, it follows from (6.64) and (6.65) that the drift losscone instability is suppressed, e.g., if a ~ 1 0 “2, Ti/MiV2 ~ 1 0 “2. In [6.26] a numerical calculation of the critical value of p ja as a function of a for various (3Xand Ti/M\vl has been performed. The plot given in that paper shows that the dependence of the critical value of p j a on (3xis unimportant when a ~ 1 0 “2.
6.5
Ioncyclotron excitation of transverse drift waves
Consider shortwavelength perturbations with kz = 0 and E B 0. According to section 2.2.2, such perturbations are described by dis persion relation (2.57) with e33 in the form of (2.58)(2.60). We restrict ourselves to the case of a Maxwellian plasma when the ex pressions for €3*3 and 6e33 are of the form of (2.61) assuming also VT = 0. In this case the /operators in (2.61) turn into differences 1 —ujn/u where u n is associated with the respective charge species. The plasma pressure is assumed to be low, (3 0 corresponds to that of rj < 0. If the usual drift effects are small compared with the current ones, i.e., Vj > V*, we have for the Maxwellian distribution
4o = (fcADi)2{ 1 + Z[(u  Ui)/kvTl}}
(7.38)
where Z is defined by (2.25).
7.6
Dispersion relation for electrostatic perturbations
The electrostatic approximation can be justified for a sufficiently small (3 and also for the case when the transverse wavelength of perturbation is small compared with the electron Larmor radius, k±pe 1. (When kj_Pe > 1, the components €02 and 603 are small due to the smallness of the Bessel functions.) We represent the dispersion relation of the electrostatic approxi mation in the form (cf. (2.80)) 1 + 4o + eoo = 0*
(7.39)
For e[,o one should substitute here the expressions given in section 7.5.2. As for in the case of Maxwellian electrons, it can be presented in a similar form to [2 ], equation (1.106): e(e) £oo —
™ ^ Jn (^ ) V l _ u ,X / { (*e + (£e 1) zeSle) u  nfie  Wde  kzvz J J ‘ L \\ (7.40) _
ri
Perturbations in a Plasma with Transverse Current
115
Here the notation of section 2 .1 .2 is used. Note also that we operate in the reference frame with electrons at rest. Note that the terms with le — 1 in (7.40) are important in the limit of highfrequency perturbations and cold electrons, i.e., for u >> (fie, k±vTe). They are similar to the terms with d2F/dxd£ in (7.35). Let us consider some particular cases of (7.40). 7.6.1
The lowfrequency approximation, with respect to electrons, for (3 < 1 In the lowfrequency approximation, only the term with n = 0 is important in the sum of (7.40). If (3 < l,V de/Vj is small, of order (3. Therefore, the term with in (7.40) must be omitted. This term was retained, in particular, in [7.7, 7.8, 6.16]. Therefore, the corresponding results of these papers are subject to inaccuracies. For small k±pe and kzVTe/u the abovementioned small term of order (3 is in the sum with formally small terms of order k\p2 e. There fore, the term with u>de must be taken into account when fc2 p2 e < f3. However, the last inequality means k± < u pe/c, and this is simply the wellknown condition of violation of the electrostatic approxima tion. A description of perturbations with k± < upe/c will be given in section 7.7. When kz = 0, an exponentially small effect of resonant interaction of electrons with perturbations is also associated with the term u>de. Such an effect was considered in [7.9]. However, for a quantitative study of this effect, it is necessary to go beyond the electrostatic approximation. Thus, the correct lowfrequency limit of (7.40) for (3 « 1 is given by 4o = (k\De)~2[1 + U ( W I
I vTe)I0(ze) exp(*e)].
(741)
Let us now discuss under which conditions the gradient effects are important in this expression. When z e > 1 o t u / \kz \ vTe < 1, the contribution of the gradient effects (they are formally associated with the departure of fe from unity) into (7.41) is of order u>*/a;, i.e., allowing for u ~ u>j, the contribution is of order V*/Vy Since (3 < 1 , we conclude (allowing for (7.23)) that the last ratio is not small compared with unity only if nB < (ftn,KT). Therefore, in situations where the characteristic lengths of the magnetic field and density gradients are of the same order of magnitude (e.g., in the case of a shock wave or a soliton), the gradient contribution into (7.41) can be neglected. Then € 00 —
(& ^ D e )
2 [1 +
%{&/ I kz
 vT e )^ o (^ e )e X p (^ e )]
(7 * 4 2)
116
Electromagnetic Instabilities in an Inhomogeneous Plasma
Dispersion relation (7.39) with 6 qo and €oo °f the form of (7.38) and (7.42) was analysed in [1 ], Chapter 13, and in the papers mentioned there. When ze kzvTe, the leading terms in (7.41) are mutually compensated. As a result, the contribution of /e—1 becomes more substantial. It then follows from (7.41) that
Dispersion relation (7.39) with 4o and defined by (7.36) and (7.43) is simply equation (2.60) of [2]. It was studied in section 2.5 of [2 ]. 7.6.2
Shortwavelength perturbations, with respect to electrons, near the electroncyclotron harmonics The region of very large wave numbers, k±pe > 1 , is of interest due to the fact that excitation of perturbations with frequencies u> > fie can be related to such k± when Vj u> and if one allows for the above inequalities. When k±pe > 1 , the components £ 02 and eQ3 are small for ar bitrary (3 (see (1.64)). Consequently, in this case the dispersion relation of the electrostatic approximation (7.39) has no restrictions with respect to (3. For the given expression (7.40) for 1S simplified. In the present section we consider simplification of €qq assuming (3 de or kzvz are comparable to or larger than the electron
Perturbations in a Plasma with Transverse Current
cyclotron frequency and (7.37))
117
(cf. section 2 .2 .1 ). We then obtain (cf. (2.52)
Cm = (fcADe)2[l + iDW1/2ie(kVTe)~1]
(7.45)
In this case the electrons behave in perturbations as unmagnetized particles. The dependence of the righthand side of (7.45) on the magnetic field (through the operator /e) is due to the fact that this field is in the expression for the equilibrium distribution function. We shall use equation (7.45) in section 8 .6 . 7.6.4 The highfrequency approximation for cold electrons In the case of cold electrons and for u > fie it follows from (7.40) that (cf. (7.36)) 4o = Wpe/W2.
(7.46)
Note that in this approximation the gradient term cancels out from the expression for €qo • Such a cancelling out is formally due to the compensating additional term with / — 1 in (7.40). The highfrequency approximation for cold electrons is important, in particular, in the problem of Buneman’s instability, see section 1.6 of [1 ]. The authors of [7.10] discovered the role of the density gradient in this problem. They started with the expression for e0o of the form of (7.40) but without the compensating term with Ze — 1. As a result, they have arrived at the conclusion that the righthand side of (7.46) contains the same gradient term as (7.43). Hence, it is clear that the analysis of the role of the density gradient in the Buneman instability in [7.10] is erroneous.
7.7
Description of electromagnetic smallscale perturbations of lowfrequency, w ith respect to electrons
In studying the electromagnetic perturbations with frequencies smaller than the electron cyclotron frequency, it is convenient to use ea/?(a,/3 = 1,2,3) as defined in section 1.1.4, instead of €ik(ijk = 0,2,3). The starting dispersion relation is then written in the form
of (1.1). 7.7.1 General expressions for eap (a ,(3 = 1,2,3) The general approach for obtaining €ap(a,/3 = 1,2,3) from was given in section 1.1.4. Following this approach, we find that in the
118
Electromagnetic Instabilities in an Inhomogeneous Plasma
case of a plasma with a transverse current r ft (1) n + €'(112 ) en = e(0) 6n _L , 6(1) £12 — £(0)
e 12
(7.47)
+ e 12
13
Here e(0) + £(i) — 1 1 h t 00 + 6n € 12
pi +
u j,
k2uQ,e
^de KBky
UJ
1
uj

UJ,pi +
(7.48)
u j }
UJ
When k±pe appear in the righthand side of (7.48). Thereby, we have extracted in explicit form both modifying factors associated with the transverse current: first, a shift of the ion component relative to the electron one which manifests as a modification of (see section 7.5.2), and, second, an additional term with ujj/uj which is a consequence of the new equilibrium condition (7.19). 7.7.2 Description of perturbations with kz = 0 When kz = 0, the dispersion relation reduces to (2.35) with en and e12 of the form of (7.47) and 622 of the form of (2.40). In this case the expressions for and are given by equations (2 .3 7 ) and (2.39). For 4o one °f the expressions (7.36)(7.38) should be substituted in (7.47). In particular, when uj < kv?*, in the zeroth approximation in k±pe and when neglecting small imaginary terms of order u/kvru the following dispersion relation is obtained (cf. (6 .1 ))
119
Perturbations in a PlsLsma with Transverse Current
We shall use equation (7.49) in section 8.3. 7.7.3 Description of perturbations in a plasma with cold electrons When.Te —► 0, in neglecting the transverse electron inertia, only the values with superscript zero remain in the righthand sides of equations (7.47). In the limit Te —> 0 one should use the identity ^De/^be = in the second equality of (7.47). The compo nent €33 is defined by €33 = Wpe/w2 (7.50) while the components e23 and e22 as well as e13 vanish. In the limit 633 — *■0 0 the dispersion relation reduces to in  N 2cos20 = 0.
(7.51)
en = €n}  4 f / i V 2.
(7.52)
Here Substituting
I
and
from (7.48) into (7.52), we find
— r(0  UpeKn^y ,
Wpe
(,
(7
^ p i+ ^ A
—
—
)■
(753)
In the case of cold ions, i.e., when —►0 and 633 is of the form of (7.36), using (7.52) and (7.53) we arrive at the dispersion relation k2k2c4Q2
k2c\u2
Wpe
(WWj)2
= 0.
(7.54)
Initially, this result has been obtained in [7.11] and has then been reproduced in [7.4, 7.12]. Note that in the absence of a transverse current, i.e., when Vj = 0, it follows from (7.50) that u 2 = k2c\+ k2k2 zc40,2 J u j^ .
, (7.55)
For k2 = 0 this corresponds to fast magnetoacoustic waves with the dispersion law ui2 = k2c\ (7.56) while if kz is sufficiently large, equation (7.55) reduces to the disper sion relation for whistlers, u 2 = fc2fc2c4ft2/ u £ = w2.
(7.57)
120
Electromagnetic Instabilities in an Inhomogeneous Plasma
Allowing for the transverse electron inertia and for terms of order I / 633 , we find, instead of (7.51), ( 6 n + ^6 n ) ( l — iV2 /e33 ) —N 2 cos2 9 ~ 0
(7.58)
8en = u l J n l
(7.59)
where For the case of cold ions, it follows from (7.58) and (7.59) that
t
UJ
KJ3 /
\
[UJ 
UJj ) 2 J
uj2
k2
(7.60) Here q is the parameter characterizing the electromagnetic character of perturbations and is defined by ? = Wpe/c2fc2
(761)
and u>lh = is the square of the lowerhybrid frequency. Initially, equation (7.60) has been obtained in [7.4] (see also [7.12]). Evidently, when q > 1 , equation (7.60) reduces to (7.54). On the other hand, when q j/u>)(l — Kn/^e). Therefore, the corresponding results of [7.13] are wrong. 7.7./
Dispersion relation for perturbations with k±pe cA than for Vj < cA. However, one should allow for the assumptions Vj < vTe and Ca > (Me/M;)1/2^ . Therefore, 7max < ( M / A Q ^ W
(8.5)
In allowing for the density gradient, k„ 0, in (7.54) it proves that an additional destabilization is associated with this gradient if d in n 0/d in B 0 > 0
(8.6)
and some stabilization takes place in the opposite inequality. Con dition (8.6) can also be presented in the form KnVj/Sle < 0
(8.7)
which coincides with equation (2.65) of [2]. Note that inequality (8.7) is a necessary condition for an electro static instability of perturbations with kz = 0 (see section 2.5 of [2]). However, it is clear from the above that, due to electromagnetic effects, perturbations with kz = 0 can be unstable even under the condition opposite to (8.7).
128
Electromagnetic Instabilities in an Inhomogeneous Plasma
The results presented here have been obtained mainly in [7.4], see also [7.11, 7.12]. The instability of a plasma with Vj > eA discussed here can be called the electromagnetic modified twostream instability (see, e.g., [8.4]). The authors of [7.13] have arrived at an erroneous conclusion which states that the modified twostream instability is impossible when Vj > cA. This erroneous conclusion has been reproduced in [8.58.8] and some other papers. In section 7.7.3 we have elucidated the reason for this error: it is related to the fact that the authors of [7.13] have used the approximation of a homogeneous plasma and ho mogeneous magnetic field, even though one should take into account equation (7.22). A possibility for the instability for V} > cA has been pointed out in [7.4, 7.12, 8.9, 8.10]. Certain applied topics, where the notions of the modified twostream instability are used, were mentioned in the introductions of Chapters 7 and 8. In addition, we note the problem of interaction between a penetrating neutral gas and a magnetized plasma [8.118.15].
8.2
Electromagnetic currentprofile instability
Consider some consequences of (7.72)(7.77). We assume for simplic ity that Ve ^ 0 only in a certain region with the characteristic size 6 smaller than the wavelength of perturbations considered, k6 I
(8.10)
Transversecurrentdriven Instabilities
129
For a significant variation of the density and the magnetic field, i.e., for Kn ~ k b — l/£ , the growth rate is of order 7
~ (c/ (M e/M 0 1/4. (8.17) For smaller values of Vj the instability should be considered as an ioncyclotron one; this is meaningful in the case of a rotating plasma. Such a situation will be presented in section 8 .7 . 8.3.3 Stabilization in a hotelectron plasma with f3 >> 1 Assume (3 > 1 ,Te > Tn and u d > (un, W j ) . Then, neglecting resonant electrons, by means of (7.49) we find (cf. (6.59)) w = 0J*i + Wj
(8.18)
In allowing for resonant electrons, we assume VT = 0. Then we obtain for Wj < w*e
7 = 2nuf//3 wDe 
(8.19)
Transversecurrentdriven Instabilities
131
It can be seen that, as in the case of the drift losscone oscil lation branch (see section 6.4.2), the perturbations considered are damped due to electrons. The ion resonance is unimportant since its contribution in (8.19) is small, of order p \ Kb (cf. section 6.4.2). The authors of [8.16] have studied the effect of the magnetic field gradient on the lowerhybriddrift instability in a plasma with a transverse current using the electrostatic approximation. However, such an approximation is invalid in this problem, so that the results of [8.16] appear doubtful.
8.4
Electrostatic and electromagnetic electronacoustic instabilities
Consider the perturbations described by equation (7.58) for kz ^ 0. When £qq is of the form of (7.37), this equation reduces to k2 zT{
1 — + ^ ( 1 uj 2
UJ
+ *lPei
(8 .20 )
1/2
17T
+
UJ — UJ\ — UJn
12
= 0.
k ±vTi
For VT = 0 this result was originally obtained in [7.5]. When (3 < 1 and q < 1, we find from (8.20) the dispersion relation of the electrostatic approximation 1
+ k l p2 UJ
k2 zT\ iTT1/2 2 ■+
M \ u j2
uj
—
u j }
k±V Ti
— UJn
i ?
= 0. (8.21)
In the approximation of a homogeneous plasma, equation (8.21) describes the current excitation of the electronacoustic oscillation branches. Such an excitation has been considered in section 13.2 of [1]. It was originally studied in [8.17]. The gradient terms in (8.21) modify the electronacoustic oscil lation branches. When in (8.21) one takes into account also the effects of finite electron temperature, the modified electronacoustic branches become hydrodynamically unstable (see section 3.2 of [2]). According to (8.21), when uj} = 0, the kinetic gradientdriven electronacoustic instability is also possible, which is an extension of the kinetic instability considered in section 2.1.3 of [2] into the region kz ^ 0. Neglecting the usual gradient terms in (8.21), i.e., assuming u n\—> 0, we have
1/2
17T
1 + k2 ±p i 
u M\u>2 + k±vT l (u ~ wi) _ O’
( 8 .22 )
132
Electromagnetic Instabilities in an Inhomogeneous Plasma
This equation describes the electrostatic currentdriven electronacoustic instability. Using (8.22), one can find that condition (8.17) is also necessary for this instability. Then k±pei ~ 1. In the essentially electromagnetic limit, q > 1, equation (8.20) reduces to u
2\
J
uj
u2
Lu,w„i(V1  22/J )l=o.
k±VT\l
(8.23) Qualitatively, it hence follows that the results of the electrostatic approximation hold up to j3 ~ 1. A quantitative analysis of equation (8.23) as well as of the more general dispersion relation (8.20) has been performed in [7.5].
8.5
Oblique currentdriven instabilities with k±pe < 1 and w
j > kvn. In this case Cqo is of the form of (7.36). Then equation (8.24) describes the shifted highfrequency ionacoustic oscillations with the dispersion law (u  u>j)2 = k2Te/Mi = k2c2 s
(8.25)
where cs is the ion sound velocity. As is wellknown, this result is true only if Te > T\. In order to consider the instabilities associated with oscillation branches (8.25), the electron Landau damping must be taken into account, i.e., terms of order u /k zvTe. Then we arrive at the disper sion relation M[(u —u>j)2
\kz \vTel
neV
2/J
v
;
For Vj ™ 0 this equation describes the gradientdriven highfrequency ionacoustic instability discussed in section 2.2 of [2]. For instability it is necessary that p j a > (.Me/Mi)1/2. (8.27)
133
Transversecurrentdriven Instabilities
In neglecting the gradient terms, equation (8.26) describes the currentdriven highfrequency ionacoustic instability if ^ > cs.
(8.28)
Such a type of instability, with the exception for ze > 1, was dis cussed in section 13.3 of [1]. The approximation 2e 1 in section 8.6. 8.5.2 Electromagnetic perturbations with u > k±VT\ For simplicity we assume nn = nT = 0. Then it follows from (7.67) that /, ^(w
v2
Wj)
k2Te\ ( ,  — J (w (« 
,
k2ky & i\ —  J
W j ) 
u j
22r2 n cAk = 0. (8.30)
Comparing (8.30) with (8.25), it can be seen that, when electro magnetic effects are taken into account, we have, in addition to the highfrequency ionacoustic oscillation branches, still two oscillation branches which, in connection with (7.57), can be called the modi fied whistlers. Note also that, when (u>,u;j) > _cs, equation (8.30) reduces to (7.54) with Kn = 0. In addition, it should be noted that, in the limit of small kz and Vj = 0, equation (8.30) yields W2 = (c2 + cl)k\
(8.31)
This corresponds to the fast magnetoacoustic waves in a plasma with f3e ~ 1. It is significant that, in contrast to (8.24), equation (8.30) has solu tions with u j TJ, so that the starting assumption u j ^ k±v?\ is valid). For such solutions, equation (8.30) reduces to (cf. (6.78))
+
= o.
(8.32)
The oscillation branches with uj the ion growth rate is exponentially small. Therefore, the analysis of the case  u j —ujj < k±vTl is of interest. This is the goal of section 8.5.3. 8.5.3 Electromagnetic perturbations with uj 7 , i.e. for (cf. (6.29)) 0 > ( * Y ' 4________l________ ~ V Mi y
n1/2(l + k 2 A jy 3/4 •
(8 44)
1
J
137
Transversecurrentdriven Instahilities
According to (8.36), in this case equation (8.40) is replaced by the following (see also (7.44)) 1 + k2A2e  , j , h2Te
Mi(u  Wj) 2
 K r\ = °.
7T \ f e ( ^  n Q e 
o>de) /
(8.45)
Then, similarly to (6.34), we find that the growth rate under condi tion (8.44) is of order Qe I
/ „ UJ \ !/3
1 +T2AL This result has been obtained in [8.3]. Note that, when both sides of inequality (8.44) are of the same order, equation (8.46) means qualitatively the same as (8.43) (cf. the analogous situation in section 6 .2 ).
8.7
Instabilities in a rotating plasma
When 7 > fii or u>di > fii, the theory of instabilities in a rotating plasma reduces to that given above. In studying instabilities with (7 >k>di) < one should modify the expression for €qo* Then we find that has a form similar to (7.40) but with the substitution u

00 ~ A;2 A2i
1 _ W
T2f/• sV  Uj  ^ n i { l + r i [ ( M j v l / 2Ti)  1] } \ \
2^\ J ^ )
wwjnfijwdi
/)'
(8.47) Here we have omitted the term with I — 1 which is unimportant for ioncyclotron instabilities. Substituting (8.47) into respective dispersion relations, one can consider a great number of varieties of the ioncyclotron instabilities. However, we shall dwell only upon one of these, namely, upon the cyclotron modification of relations (8.16) when condition (8.17) is violated. In this case the starting dispersion relation has the form
£i(u>u>j rafy  w di)
Here we have neglected the density and temperature gradients and have assumed the electrons to be cold. In contrast to section 8.3.2
138
Electromagnetic Instabilities in an Inhomogeneous Plasma
but by analogy with section 6.3.1, we have allowed for the terms of order k2p2{. Note that equation (8.48) is similar to (8.45). In addition, there is an analogy of (8.48) with (6.28). Therefore, we shall give results following from (8.48) without additional calculations. It is necessary for instability that
Vj/vn > (M e/M 0 1/2.
(8.49)
The real part of the oscillation frequency is approximately given by the first expression of (8.16). The term in (8.48) can be neglected when /3 < (M e/M i)1/2. (8.50) In this case 7
~ /31/ 2 (M e/M i) 1/4 fti.
(8.51)
When a condition opposite to (8.50) takes place, i.e., when /3 > (Me/M 0 l/2
(8.52)
equation (8.51) is replaced by 7
~ (/?Me/M i) 1/3 fli.
(8.53)
This growth rate is smaller than (6.34) which is explained by the smallness of the frequency of (8.16).
9
Hydrodynamic Description of Collisional Finite/? Plasma Perturbations
Instabilities in an inhomogeneous finite/? plasma can be studied on the basis of two approaches: (i) kinetic, and (ii) hydrodynamic. The kinetic approach has been given in Chapter 1. Now we consider the hydrodynamic approach. As is wellknown, in the hydrodynamic approach, the kinetic equa tion for each particle species reduces to an equation set for some first moments of the velocity distribution function of these particles. Such a procedure is usually due to two reasons: first, to allow for collisions between particles, and, second, to obtain a clearer picture of plasma processes. These reasons stimulated a great number of approaches to the simplification of the kinetic equation. This resulted in an equation set of twofluid magnetohydrodynamics containing density, velocity, pressure and some other moments of the distribution func tion. The authors of [9.1] have analysed a suitable subset for the problem of instabilities in an inhomogeneous finite/? plasma, and have improved some of the most familiar hydrodynamic approaches allowing for the terms essential for this problem (see in detail below). The same viewpoint forms the basis of the present chapter. As will be seen below, for a complete description of the instabilities in a collisional finite(3 plasma it is necessary to deal with a rather large number of moments of the distribution function. Therefore, in the general statement of the problem, one of the main advantages of the hydrodynamic description, i.e., clarity, is in fact lost. In order to keep some element of simplicity of the hydrodynamics we be gin our presentation by obtaining the simplest set of hydrodynamic equations suitable for studying the perturbations with kz = 0 by neglecting dissipative, inertial and magnetoviscous effects, section 139
140
Electromagnetic Instabilities in an Inhomogeneous Pleusma
9.1. Such equations are found by using the drift kinetic equation. They are the basis of the socalled drift hydrodynamics of a finite/? plasma. In contrast to section 9.1, the goal of sections 9.2 and 9.3 is to find the starting equations for the complete description o f the in stabilities in a finite/3 plasma. Note that the problem of choosing transport equations adequately describing these instabilities is far from trivial. This problem has been proceeded by a simpler prob lem of describing the instabilities of a collisional plasma with zero/? which has been considered in [9.2]. Before paper [9.2] it was gener ally accepted that, in order to describe completely the instabilities of a collisional plasma , one could use the wellknown Braginskii equa tions [9.3, 9.4] or the similar Kaufman equations [9.5]. The author of [9.2] has shown that, in fact, some problems of the theory of colli sional plasma instabilities can be analysed by means of equations of the type of [9.39.5]. At the same time, according to [9.2], in some cases the use of those equations leads to incorrect results. This is explained by the fact that the totality of small parameters adopted in deriving equations of type [9.39.5] does not correspond to that adopted in the theory of gradient (drift) instabilities. Such a non conformity has been discussed in Chapter 4 of [2]; therefore, we shall not consider it in detail here. However, we should mention the fact that in the theories of [9.39.5] the transverse velocities of the plasma components are assumed to be of zeroth order, while the transverse heat fluxes are of first order. As for the ordering adopted in the driftwave theory, the transverse velocities and the transverse heat fluxes are assumed to be of the same order. This fact results, in par ticular, in difficulties in the use of the expressions for the viscosity tensor given in [9.3, 9.4]. In the abovementioned papers the viscosity tensor is expressed in terms of the velocity derivatives and does not contain the deriva tives of the heat flux, which is obviously unacceptable for the driftwave theory. These difficulties can be overcome if, in deriving the transport equations, the transverse velocities and the transverse heat fluxes are assumed to be of the same order. Such an ordering can be made by two approaches: first, by assuming the both quantities to be of first order, and, second, by considering them to be of zeroth order. The first approach has been initially developed in [9.2], where the transport equations suitable for describing the gradient instabil ities in a zero(3 plasma in a homogeneous magnetic field have been obtained, and then generalized in [9.6] to the case of arbitrary (3 and curvilinear magnetic fields. The second approach goes back to the idea used by Grad [9.7] of a 13moment set o f transport equations which has been realized for the case of a plasma in a magnetic field
Hydrodynamic Description of Collisional Finiteft Perturbations
141
in [9.8]. Recall that in G rad’s approach the zeroth order values are den sity, temperature and velocity vector (i.e., 5 quantities), as well as the heat flux vector (3 quantities) and the viscosity tensor (5 quan tities since the viscosity tensor is symmetric and of zeroth spur). At the same time, it is clear that, generally speaking, there is no rea son, to adopt Grad’s equations as being a priori acceptable for the problem of gradient instabilities, since there is no guarantee that in some cases other moments of the distribution function will not be essential. Moreover, Grad’s equations fail to reproduce Braginskii’s results. Therefore, generally speaking, the approach of [9.2, 9.6] is more consistent than that of [9.7, 9.8] for the driftwave theory. However, Grad’s approach is more obvious since, in this approach, all the moments of the distribution function have a simple physical meaning. The authors of [9.1, 9.9] have extended Grad’s approach by allow ing for some additional moments of the distribution function. They have augmented Grad’s 13 moments by 10 new moments. The 8 new moments are a tensor of the second rank similar to the viscosity ten sor and a vector similar to the heat flux. Allowing for these tensor and vector moments is necessary in order to reproduce Braginskii’s results. In addition, according to [9.1, 9.9], one should take into account two scalar moments which are not present in either [9.39.5] or [9.7] but arise in the approach of [9.2]. The abovementioned multimoment equation set is given in section 9.2. When the cyclotron frequency is large compared with the colli sion frequency, the multimoment equation set of section 9.2 is signif icantly simplified. Then this set reduces to a form similar to Bragin skii’s equations but with modified expressions for the viscosity tensor and the heat flux. Such a reduction is presented in section 9.3. The equations of section 9.3 show that, in a number of problems, besides the abovementioned influence o f the heat conductivity on viscosity , such a crossed effect as the influence o f the viscosity on heat conductivity is also significant (this effect is not described by Braginskii’s and Kaufman’s equations). In addition, the equations of section 9.3 allow for, more correctly than Braginskii’s equations, the socalled gyrorelaxation effect , i.e., a collisional thermal redistri bution among the different degrees of freedom. This effect was orig inally considered in [9.10, 9.11] in connection with the problem of plasma heating by an oscillating magnetic field. The gyrorelaxation effect is also important for the problem of collisional finite/? plasma instabilities which will be discussed below. The simplest set of hy drodynamic equations allowing for effects of order including the gyrorelaxation effect, is given in section 9.4 (yx is the ion collision
142
Electromagnetic Instabilities in an Inhomogeneous Plasma
frequency ). In accordance with the above, the equations of section 9.4 have been found by simplifying the general set of transport equations de rived by means of the moment method from Boltzmann’s kinetic equation. In section 9.5 we show that precisely the same equations follow from the drift kinetic equation with a collisional term. This, on the one hand, allows one to treat the corresponding hydrody namic equations in terms of “drift” notions , and, on the other hand, is a guarantee of the correctness of calculations made under a sim plification of the general transport equations. The simplified hydrodynamic equations of sections 9.1, 9.4 and 9.5 do not allow for inertia and magnetoviscosity effects , nor for the transverse heat conductivity and the transverse viscosity, and a series of other effects important in some problems on instabilities. These effects are discussed in the remaining sections of this chapter. In section 9.6 we explain the specificity of allowing for inertial and magnetoviscosity effects , while in section 9.7 we treat the transverse heat conductivity and the transverse viscosity. Some other collisional effects will be considered in Chapter 10 when deriving the dispersion relations for separate types of perturbations. Note also that the hydrodynamic description is sometimes used even when neglecting collisions. The approach of [9.12] called the cgl approach is an example of such a description. The c g l equations have been augmented by the magnetic viscosity in [9.13] (see also [9.14]). Such modified c g l equations have been used in [9.13] in the problem of inertial perturbations. However, the procedure of [9.13] is, generally speaking, risky since, in the problem on drift waves in a collisionless finite/? plasma, there is no small parameter, allowing one to reduce the kinetic equation to a moment equation set.
9.1
Drift hydrodynamics of a finite/? plasma
As in the previous chapters, we consider a plasma to be in an equi librium magnetic field B 0 z. In this section we shall find starting equations for describing perturbations with kz = 0 and Ez = 0. In contrast to the previous chapters, we now assume that the collision frequency, v , of each particle species is large compared with the per turbation frequency, v > u. We start with Boltzmann’s kinetic equation which, for the abovementioned assumptions, has the form W + „x . v / + £ , . 8 £ = n W + c .
(M)
Hydrodynamic Description of Collisional Finite/3 Perturbations
143
Here / is the total distribution function, i.e., the sum of the equilib rium and perturbed distribution functions, and C is the collisional term . The quantity fi is the cyclotron frequency defined by the total magnetic field which, in contrast to Chapter 1, is now designated by 2?, so that B = Bo + B y where B is the perturbed magnetic field (cf. sections 6.6.2 and 7.8), and a is the angle in particle transverse velocity space (cf. section 1.2.4). For a sufficiently strong magnetic field when > */, equation (9.1) can be solved by expansion in series in the small parameter 1/fi. A corresponding procedure for the case of a homogeneous magnetic field and electrostatic electric field was presented in Chapter 4 of [2]. Generalizing this procedure for the case involved, we arrive at the following equation
% + Ve •v /  S±—^~V •VE + £je, •V / x V ^ = C.
(9.2)
Here / is the part of / independent of the angle a , C is the average of C over a , Ve = c E x B j B 2 is the particle drift velocity in the crossed fields E and B (the crossedfield drift velocity), and = v2 /2. Equation (9.2) is simply the drift kinetic equation with the
collisional term. Assuming the function / in the lefthand side of (9.2) to be Maxwellian, we take the moments of (9.2). Neglecting the dissipative effects which will be considered below, we find the equations for the total (in the abovementioned meaning) density n and temperature T of each species:
3n
_
1 v + VE .V n + n V V E + e ,. V  g : X V  = 0
% + V v Vp + 7o;pV •F E + 7oe, •
x
= 0
(9.3)
(9.4)
where 70 = 5 /3 is the adiabatic exponent and p = nT.
9.2
Multimoment transport equation set
Equations (9.3) and (9.4) allow one to study only the particular case of perturbations of a collisional finite/3 plasma neglecting all the dissipative effects. We shall now derive equations suitable for a general analysis of the problem of instabilities in such a plasma.
144
Electromagnetic Instabilities in an Inhomogeneous Plasma
We start with the fact that sufficiently frequent collisions between particles lead to a velocity distribution close to the Maxwellian case, / = /o (l + * )
(9.5)
/o = n(M /2w T )3/2 exp [M (v  V f/ 2 T ]
(9.6)
where as long as the value of $ is small, $ < 1. The quantities n ,V and T are the density, velocity and temperature of the corresponding plasma component, respectively, if $ satisfies the condition
J ( v  V ) kf 0$ d v
k = 0 ,1 ,2 .
(9.7)
As a function of the “random” velocity vector w = v — V , $ can be written as a sum of scalar, vector and tensor parts $ = $0 + Wi$i + (wiwk  6ikw2/3)$ik + . . .
(9.8)
The ellipsis in the righthand side represents tensors of the third and higher ranks which are neglected. A general expression for the scalar $ 0 satisfying the first and third conditions of (9.7) can be presented as a series of the SonineLaguerre polynomials L ^ 2\x), where x = Mw2/2 T , starting with 1 = 2:
oo
a^L\m (x).
$0=
(9.9)
1=2
Similarly, the vector $,• satisfies condition (9.7) if OO
= Y / 4 ,)L (j3/2\x). /=1 Since we impose no restrictions on the tensor as a full series oo
=
(9.10) it can be written
(9.11) 1=0
Series of the same type can be found for the tensors of higher rank omitted on the righthand side of (9.8). The coefficients a ^ ,a ^ and af2 are higher moments of the dis tribution function. The quantities a,1^ and have the simplest
Hydrodynamic Description of Collisional Finitefi Perturbations
145
meaning. They are related to the heat flux q and the viscosity ten sor irik by
lw
q‘
“a =
( 9  1 2 )
(9 13)
This can be proved by means of (9.5), (9.10) and (9.11) and the definitions of q and irik:
q =
J w 3f
d«
7rik = M j ( w i w k  6ikw2/Z )f du.
(914) (9.15)
Allowing for (9.12) and (9.13), we shall also introduce the quan tities q* and ir*k determined by the relations
&
=Ijf*
““ = W * * '
(916) ( 9 '1 7 )
Using Boltzmann’s kinetic equation
dj_ _ . e / _ v x B\ df „ at + v  V f + u { B + — ) ^ = c
(9.18)
a series of equations can be obtained relating the quantities etc. This series can be truncated in different n, ways neglecting quantities with a sufficiently high number of the tensor indices. The different ways of truncating the series of equa tions lead to a closed transport equation set with different degree of accuracy suitable for describing specific phenomena. The examples of such a transport equation set suitable for studying the electrostatic instabilities in a lowfi plasma were given in Chapter 4 of [2]. To obtain the correct limiting transition to Braginskii’s results, it is necessary to take into account not only the values q and 7r (i.e., and a ^ ) , but also a and This means that in expressions (9.10) and (9.11) for and the first two terms of the series must be considered ( approximation o f two polynomials).
146
Electromagnetic Instabilities in an Inhomogeneous Plasma
In addition, the contribution of the terms with the scalar $ 0 (see (9.9)) and the tensor terms of third and higher ranks must be es timated. It appears necessary to take $ 0 into account. It is also necessary in expression (9.9) to consider two polynomials to describe correctly the quantitative effects related to $ 0. As a consequence, our problem consists in the construction of the 23 equations for n, V ,T , a\l\ a^\\ a^\a\2\ a^2\ a^3\ The first five equations of this set, obtained by integrating equa tion (9.18) with the weights 1, v, (v —V )2, have the wellknown form d n /d t + V {n V ) = 0 (9.19)
dV 1 M n —— = —Vp — V •7r + en (E +  V x B ) + R dt c
(9.20)
^ + ^ v ' v = _ v ' 9 _ (7 5 " v ) ‘ v + q (921) Here d/d t = d /d t + V •V; R is the momentum transfer (the fric tional fo rce ) R = M j ( v  V)Cdv
(9 .2 2 )
and Q is the heat transfer , Q = Y j ( v  VfCdv. In the approximation V\ tensors 7tA;i and tt^, are: (^Ap +
(Z
Ui
(d/d/, V •V ) the equations for the
2 •c
, 205 , \
(9 .2 3 )
x flj
35
—
7 rir(2)
^
48'?r^J  T C
gPi^A/?
(9.24)
‘ >A/i = ~ 6 Pi A"
Here W $ = 2 < V •V » A„ + £  < V • q » A„
OPi
Pi
J i
+ V  q * + q* V ln T i » A„
F = e E /M + V x Q  d V / d t f I = e B /M c XB
dv^
dy
dx
dV™ _ cHj°> 2fii dx dy
2 /a«g»
dq^ \
5pi V dy
dx J
2 f d q jQ> _ dq  Wg> + = *„ = ^
(W%  Wg>)]
(9.60)
(W«>  W) 
Here, as follows from (9.25),
v v  H
>
As in deriving (9.36), we have taken into account that F\ =
(9.61)
= V pJM [n. Expressions (9.60) reduce to those derived by Braginskii [9.3, 9.4] when the terms with are omitted. However, such an approxi mation is not justified in the problems considered since q± and V± are values of the same order: they are proportional to the spatial derivatives o f the temperature and the pressure. Respectively, expressions (9.59) and (9.60) differ from (9.33) and (9.57) by terms of order Due to the law o f conservation o f momentum under collisions between the ions, the terms of order uJQi do not appear in the expression for the velocity, so that one can take VI = V ^ in calculating the viscosity tensor, where v [ 0^ is given by (9.32). The same accuracy is sufficient for substituting VI into the inertial term of (9.49) since it even contains the small factor of order u.
10 Starting Equations for Basic Types of Hydrodynamic Perturbations
Using the results of Chapter 9, in the present chapter we shall find the starting equations for a number of specific types of hydrodynamic perturbations. Thereby we shall complete the preparation of every thing required for the analysis of the stability of such perturbations which will then be made in the successive chapters. In the present chapter we consider mainly smallscale perturba tions. As in Chapter 1, we assume that they can be described within the scope of the local approach of the WKB approximation. Accord ingly, our goal will be the derivation of the local dispersion relations for such perturbations. In addition, a part of the present chapter is devoted to largescale perturbations , which we shall discuss in detail below. According to Chapter 2, an important property of lowfrequency longwavelength perturbations of a collisionless plasma is the split ting of their dispersion relation into two dispersion relations, one of which corresponds to the Alfven perturbations (inertial), while the second corresponds to magnetoacoustic and other perturbations (also called inertialess). This fact is important for both the physical viewpoint (to understand the nature of instabilities) and the formal viewpoint, since it substantially facilitates the analysis of the pertur bations. Therefore it appears to be important to find out whether the splitting takes place in a collisional plasma. The study of this problem is one of the objectives of the present chapter. As a result, we show that the splitting does take place. In addition, our analysis allows one to give a hydrodynamic interpretation of the splitting and justifies the use of the terms inertial and inertialess perturbations. Our presentation starts with the derivation of the dispersion rela 156
Equations for Basic Types of Hydrodynamic Perturbations
157
tions for the simplest case of perturbations with kz — 0, neglecting dissipative effects. This is the goal of section 10.1. In such an ap proximation the inertialess and inertial perturbations come to be mutually independent. Then, in section 10.2 we explain how the dis persion relations for the abovementioned perturbations are modified when allowing for the dissipative effects of orders ujv\, Uik]_pf/uj and vxe/u;, and find the dispersion relation for the transverse drift waves (* , = 0). Section 10.3 is devoted to deriving the dispersion relations for perturbations with kz ^ 0. There we consider the perturbations with kz ~ (veujy/2/vTe, those with kz ~ u>/vt[ in a weakly collisional highfi plasma and those in a strongly collisional plasma . Finally, in sections 10.4 and 10.5 we deal with the perturbations in a plasma flow with an inhomogeneous velocity profile (the socalled shearvelocity flow). In section 10.4 we derive the dispersion relations for the inertialess perturbations in such a plasma flow. Section 10.5 concerns the inertial (Alfven) perturbations. As in sections 10.110.3, we discuss smallscale perturbations in section 10.4. In section 10.5 both smallscale and largescale perturbations are studied. A more detailed analysis of the problems considered in the present chapter can be found in [2.1, 2.2, 9.1, 9.9, 10.110.6]. The results of these papers are used in our presentation. The fact mentioned above, namely, the splitting o f the oscilla tion branches in a collisional plasma into inertial and inertialess ones, is presented in section 10.1 and in some subsequent sections. Our analysis reveals a hydrodynamic feature of this effect: the set of hydrodynamic equations, obtained by neglecting the inertial and magnetoviscous effects, allows one to obtain a dispersion relation in spite of the fact that the number of hydrodynamic variables in cluded in these equations (density, temperature, etc.) exceeds the number of equations. In this connection, the procedure of deriving the hydrodynamic dispersion relation is based upon the transition to the socalled “canonical variables”, the number of which coin cides with the number of equations. Such a procedure for deriving the dispersion relation was initially used in [2.1, 2.2]. The dispersion relation for inertialess perturbations given in sec tion 10.1, was initially obtained in [2.1] in the approximation of straight magnetic field lines. As will be explained in Chapter 15, the same dispersion relation is obtained in the case when transverse inertia and magnetic viscosity are small compared with the effects due to the magneticfield curvature. For this case the dispersion re lation for the inertialess perturbations has been found in [3.1], where such perturbations have been termed entropy waves. Following [3.1], we use the term “magnetodrift entropy perturbations” for the iner
158
Electromagnetic Instabilities in an Inhomogeneous Plasma
tialess perturbations with kz — 0. The authors of [10.7] studied the nonlinear regime o f the entropy waves. They have shown that the set of nonlinear equations for the entropy waves, as well as that of linear equations, is autonomic. Autonomy means that, in analysing these equations, the additional (i.e., inertial) equation is not required since, similarly to section 10.1, one may express the hydrodynamic variables in terms of a smaller number of canonical variables. One more remark should be made concerning the perturbations in a collisional plasma flow with an inhomogeneous velocity profile (sections 10.4 and 10.5). Note that the dispersion relations for the collisional plasma flow were initially obtained in [10.8]. However, the results of [10.8] are inaccurate since, firstly, in the case of the inertialess perturbations, the viscosity tensor has been incorrectly linearized, and, second, in the case of Alfven perturbations, this tensor has been taken in an inadequate form.
10.1
Dispersion relations for perturbations with kz = 0, neglecting dissipative effects
10.1.1 Inertialess perturbations Linearizing (9.3) and (9.4), we find
iu>— + n0 •P /
po ^
+ ^ ( k „  kb ) +
±>q
\
.
n no
CE V/ Jj 0
no
 w*) = 0 (10.1) .
 i — [ u  2 j 0ui d )  1 7 0 wD +  5HKp  7oKb)
Po
+ h o% {u  2u>* + u n) = 0.
(10.2)
DO
Here the subscript zero denotes the equilibrium quantities, while a tilde denotes perturbations, kp = d ln p /d x . The remaining nota tions are clear from the previous chapters. Both equations (10.1) and (10.2) are associated with each particle species. For simplicity we omit the subscripts of the particle species. From (10.1) and allowing for the quasineutrality condition ne = n, we obtain 4w(pi + pe) + B ,B o  0 . (10.3) This is the socalled perturbed pressurebalance equation. By means of (10.3) we exclude the field B z from (10.1) and (10.2). Following [3.1, 2.1] we introduce the variables X ,Y ,Z ( canonical
Equations for Basic Types of Hydrodynamic Perturbations
159
variables ) defined by the relations X = (l + \
Toe/
u
P i+ P e
Kp P0i + POe
\Poi icEy
POe J
(10.4)
Bo
n_ _ Kn pi + pe n0
Kp Po\ + POe
Then, from (10.2) and (10.1) we find
U>DiX + (Kg —Kn)Y —UZ = 0
u(Ti/Te)X
+ 7owDi Z = 0
(10.5)
[w(l  Te/T i)  2 7 0o;Di] X + (kp  70k b )Y = 0. Here for simplicity we have omitted the subscript zero at Te and T\. Equating the determinant of (10.5) to zero, we arrive at the dis persion relation [3.1, 2.1]
dispersion relation (2.14) examined in section 3.1. The analysis of (10.6) will be performed in section 11.1. Note that the variables X,Y, and Z defined by relations (10.4) are constructed by means of the four perturbed quantities : pi,pe, n and E y. Therefore, it is clear that the equation set (10.1)(10.3) is incom plete] it should be supplemented by one more equation connecting the abovementioned quantities. One can take equation (9.49) as the supplementary equation. Nevertheless, this incomplete equation set allows one to obtain dispersion relation (10.6). This reflects the splitting of lowfrequency longwavelength oscillation branches into two mutually independent kinds (this fact was noted in Chapter 2 and in the introduction of Chapter 9). In fact, it follows from the previous equations that the perturbations described by dispersion relation (10.6) can be studied by means of the hydrodynamic equations (9.3) and (9.4) derived by neglecting the plasma transverse inertia. This provides the basis for calling perturbations of type (10.6) inertialess, which we have done since Chapter 2. Conversely, the supplementary equation (9.49), as mentioned above, depends very substantially on the transverse inertia. This
160
Electromagnetic Instabilities in an Inhomogeneous Plasma
equation can be satisfied in the following two cases: (i) when (X , y, Z) ^ 0, i.e., under condition (10.6); in this case it will give the missing connection between the quantities contained in (10.1)(10.3); and (ii) when X = Y = Z = 0; resulting in some other dispersion relation which will only correspond to inertial perturbations. Hence, the given considerations lead to the conclusion of the ex istence of two mutually independent types o f perturbations in a col lisional plasm a , analogous to the case of a collisionless plasma.
10.1.2 Inertial perturbations Using equations (9.32), (9.33) and (9.57), we linearize (9.49). As a result, we arrive at the equation
/ „ XPi . Kv / 3 . cEv + ( uj — 3 u ;Di )  b i — ( u — T ^ D i ) ” ^   — 0 . Po
Bio
2
^pi
In terms of the canonical variables X , y, Z defined by relations (10.4) (for T0e = T0i = T) this equation implies
[u j2 — U j( u * i +
U>Di) + ^
p i ^
D
i — ^ D i(^ “
 KB(u  ^wDi)y 
3 c JD i)X
(10.8)
= o
where a is defined by (2.10). At the same time, we have equations (10.5) so that we now have the complete equation set. In section 10.1.1 we found the dispersion relation (10.6) assuming X ,y , Z to be finite. At the same time, according to the discussion above, equations (10.5) can also be satisfied for X = Y = Z = 0. In this case, under the condition
Bz ^ 0
(10.9)
the dispersion relation (4.1) is obtained from (10.8) corresponding to the Alfven perturbations. Thus, it is clear that the dispersion relation for the inertial perturbations in a collisional plasma is the same as that in the case of a collisionless one [2.1, 2.2]. Taking into account (10.3) and (10.4), from the condition X = y = Z = 0 we find that, in the inertial perturbations,
Equations for Basic Types of Hydrodynamic Perturbations
161
n _ k„ B z no
Kb Bq
JL  I I Toi
Toe
(10.10) Kj3 Bo
iu>Bz/ B 0 = KscEy/Bo where kt = d ln T /d x . The last equality in (10.10) can be repre sented in the form V •VE = 0
(10.11)
where, as in section 9.1, VE — c E x B / B 2 is the crossedfield drift velocity. Therefore, it is clear, that the Alfven perturbations in an inhomo geneous plasma are characterized by the crossedfield drift velocity corresponding to the incompressible plasma motion (see (10.11)). Due to the equilibrium magneticfield inhomogeneity such a motion leads to perturbations of the ^component of the total magnetic field, i.e., to compression and expansion of this field (see (10.9)). This is in contrast to the case of a homogeneous magnetic field where the Alfven perturbations do not result in the compression and expansion of the field lines. According to (10.10), the density, temperature and plasma pressure are perturbed together with the magnetic field.
10.2
D ispersion relations for p ertu rb ation s w ith kz = 0 allowing for dissipative effects
10.2.1 Effects of order u/v\ Let us linearize equations (9.38)(9.40) using the quasineutrality condition (cf. section 10.1). Then (10.1) and (10.2) are replaced by
162
Electromagnetic Instabilities in an Inhomogeneous Plasma
i(u> 
2 7 0u;Di)^Poi
+ ( kp  70kb)—
Bo
IT o ^ D ino
P> , cE A 1W— + Kp—
+ i7o(u,  2u£ + « ni ) j + + 6.41^ — 9 ^
Poi
Ti
Bz
j Oi
D o ,
DO /
wdi7?;— uTl— )  0 (10.13)
B, P PiX + Pe Bo 2 Poi "HPOe The quantity pix in the last equation means
(10.14)
Xzz 0.96 poi • Pi 1. K cE, —10J P .X = P 1 T  = P 1 — Poi
, .4 2 / Bz Ti B n Ud> Toi 141 \ T'~"—
p
B
(10.15)
As before, the electron density and the electron pressure are deter mined by equations ( 1 0 . 1 ) and ( 1 0 .2 ). Assuming T0e = To; for simplicity, we introduce the following vari ables similar to (10.4):
X,
=
PiX ~
Pe
Po
y ,=
"2 i± A + i 4 Kp Po DO K„ PiX + Pe *x = ^n0 Po
(10.16)
Here p0 = Poe+Poi In terms of these variables the set (10.12)(10.14) takes the form (cf. ( 1 0 .5 )) wDiXx + ( k b ~ K„)yj. — u>Z± = 0
[u> + e(A + Z))]Xx 4 (Kp/u)eAYx + (7owDi — —0 [27ou;Di + t{A + D)]X± + [kp  7 0kb + (Kp/u)eA]Yj_  eAZ± = 0. (10.17) Here .0.96 u> e = 110 A = oj — 0.59wDi (1 0 .1 8 )
v,
D = ~rr——(w + 22.02woi)41 w
Equations for Basic Types of Hydrodynamic Perturbations
163
The determinant of equation set (10.17) yields the dispersion re lation. When (3 /i/j corresponding to the longitudinal viscosity and related ef fects . When one takes into account similar terms in deriving the dis persion relation for the inertial perturbations , it can be seen that all these terms are expressed only in terms of the variables X ±,Y ±,Z ± (see (10.16)). This means that the dissipative effects of order u/v\ do not appear in the dispersion relation for the inertial perturbations. In other words, even when the terms of order ujv\ are allowed for, equation (4.1) still holds.
10.2.2 Effects o f order v\k\p\Iu Using (9.59)(9.61), we find that allowing for terms of order vxk\p?/u> modifies the ion heatbalance equation (10.2) and the inertial equa tion (10.7) which now take the form, respectively: .
Pi
/
^
.
n
.
n
/
.
\
 1 — ( u  2 7 oU>Di + 1Ol)  1— ( 7 oU>Di  lot) P oi
n°
.
(10.20)
+ ^  ( « p  7 o «b ) + iT o ir tw Oo Do
+ w „i) = 0
+^(2(“_3"Di)+l io) icEv
„
(1021)
9 . \
where a = (4 /3 )vik±pf. These equations are supplemented by the continuity equation (10.1), by the electron heatbalance equa tion (10.2) and by the perturbed pressurebalance equation (10.3)
164
Electromagnetic Instabilities in an Inhomogeneous Plasma
which are unchanged. From (10.20), (10.21), and (10.1)(10.3) after transforming to the variables X , Y, Z we find the dispersion relation [10.5]
X {2w3 + w2 [wDi(27o  3) + 2w*i]
(10.22)
■f wcjQi [3u>Di + u>„i(27o — 3) — 6w*j]
+ 3coSi(27oWDi + 4w*i  470w„i)}. Here Q{°^ and represent the lefthand sides of equations (4.1) and (10.6) (for 71 = Te), respectively, while 8Q\ and 6Q2 are of the form (10.23)
an 2 OQ
ia 2 + 7 o/3
_»(l +fn +l) +^) 3
+ wDi(7o/?  1)  7 o j3uf„
(10.24)
The dispersion relation (10.22) will be studied in section 11.1.3.
10.2.3 Allowance fo r the heat transfer Above we neglected the heat transfer Q in the heatbalance equations (9.21). According to [9.3, 9.4] Qe = Qx =  ( 3 / 2 > ien(Te  T[)
(10.25)
where vK = 2Me/M ;re, and re is the characteristic relaxation time of the electron gas. (The explicit definition of re is given in [9.3, 9.4].) When Zoe — 7oi? allowance for Q leads to the following modifi cation of dispersion relation (10.6) for the inertialess perturbations [3.1]: + ii/ie) +
1 + 2 lo — — — ^ = o. \ /
7 o^Di (
( 1 0 .2 6 )
Since Q is expressed only in terms of X , the heat transfer con tribution into the dispersion relation for the inertial perturbations vanishes. We shall examine (10.26) in section 11.1.4.
165
Equations for Basic Types of Hydrodynamic Perturbations
10. 2.4 Dispersion relation for transverse drift waves Consider the perturbations with kz = 0 and E B 0. In this case B = B oez + B ± , where B ± = (B xyB y, 0). We start with the 2 component of the electron motion equation (9.20) written in the form een0(E ,  V;eB x/ c ) + R ,e = 0. (10.27) Here V*e = cTenp/eeBo is the electron pressuregradient drift velocity and R ze is the ^component of the electron momentum transfer Re. According to [9.3, 9.4]
R e = —snVTe  i/enMe(Vje  Vjji)e0
(10.28)
where s = 0.71, Vj = V •e 0, e 0 = B / B = h ,v e is the electron colli sion frequency (ve ~ l / r e), and V = e0(e 0 ’ V ). Linearizing (10.28), we take into account e0 .V =   + 4 B x V .
oz
B0
(10.29)
Therefore, Rze =  s n 0B x(dT 0e/dx)/B0  uen0M eVze.
(10.30)
We neglect the ion longitudinal velocity. Besides, it is clear that we have neglected the electron longitudinal inertia in (10.27). We supplement equations (10.27) and (10.30) by Maxwell’s equa tions curU.Bj. = 4ireen0Vze/c ^
E z = u B x/c k y. Taking into account V •B ± = 0, from the first equation of (10.31) we find Vze =  i k 2± cBx/4ireen0ky. (10.32) As a result, the dispersion relation is obtained (cf. (3.16)) w = u>pe + su>Te  ic^xl'e/Wpe.
(10.33)
The oscillation branch (10.33) will be discussed in section 11.2.
166
Electromagnetic Instabilities in an Inhomogeneous Plasma
10.3
Starting equations for oblique perturbations
10.3.1 Perturbations with kz ~ (VevY^/vTe In contrast to sections 10.1 and 10.2, we now consider kz ^ 0 ( oblique perturbations) and allow for finite electron heat conductivity along the lines of the total magnetic field, i.e., take qe of the form [9.3, 9.4]
( 1 0 3 4 )
We neglect the ion longitudinal motion, Vzi = 0. Taking into account (10.32), we neglect the contribution of Vze in the electron equations of motion and of heat balance, (9.20) and (9.21), as being as small as k\. Therefore, the linearized electron continuity equation is the same as that for kz = 0 and is of the form of (10.1). Using the 2 component of the electron motion equation (9.20) (cf. (10.27)), we express the magnetic field perturbation B x in terms of other per turbed quantities: B x = ik>cTof ky _ P l (1 + s) + s !!_\ . eeu> \ B 0 u m p0e n0J
(10>35)
Here u> = u —w*e — suxe (cf. (10.33)). Using (10.34) and (10.35), we reduce the linearized electron heatbalance equation (9.21) to the form (cf. (10.2)) . Pe f
l—
0
w  u > „ e\
u>  2 7 0o>De + i A ;
Poe\
+
u>
—
1 oK B
J
.n (
 i—
7oWDe  i A — ^E
n0 \
w
—
+
/
2 Wpe + Wne) = 0 .
(10.36)
Here A = (2/3)3.16fc2To/Mei/e. It is assumed that Toe = Toi = To. Introducing the variables X ,Y , Z given by (10.4), we arrive at the set of three equations similar to (10.5): + («B  nn)Y  ojZ = 0 X {u  27owDi) + Y ( kp  7qk b ) + 7oWDiZ = 0 +
2^(U ~ w« e ) ) ~ ^ 2 i ) KT ^
^(“
To^De +
~ Wp e )) = 0 .
(10.37)
Equations for Basic Types of Hydrodynamic Perturbations
167
Hence the dispersion relation follows [10.1] ( 1+ M +, : ^ +
) [ , + 7^
( l . ^
M
) ]
[x +
2/(1
+ 5 )  s]
( 1+ ! * , ) + ,[ 1+ 2 £ + ra 2 (a+ 0 ]
a£(1+s, +*> _£)}.,. (10.38)
Here x = u / u ni,y = kp/k „, and A = A / w „ j. We shall examine equation (10.38) in section 11.2.
10.3.2 Perturbations with kz ~ w /vn in a weakly collisional plasma Following [2.2], we shall now consider the perturbations with kz ~ uj/ vTi. We neglect the heat transfer assuming u> > vie. This con dition corresponds to the socalled weakly collisional plasma regime. Another limiting case, u >• vie, the leading term in the electron heatbalance equa tion (9.21) is due to the heat flux related to the longitudinal electron heat conductivity. This heat flux is described by the first term of the righthand side of (10.34). Then (9.21) reduces to e0 •VTe = 0.
(10.39)
Linearizing (10.39) allowing for (10.29), we find
Te/To = iKTB x/ k zB 0.
(10.40)
In contrast to section 10.3.1, the ion longitudinal motion is now important, Vzl 0. Then Maxwell’s equation for the field f?x(10.31) contains the difference Vze —VZi instead of Vze. Equation (10.32) is also modified in a similar manner. From that equation it then follows that the difference VZ(. —Vz\is small, of the order of k\. Therefore, one can assume VZ6 = Vzi = Vz (10.41) in the electron and ion continuity equations and in the ion heatbalance equation. In allowance for Vz / 0 the term with ikzVz is added to the lefthand side of equation (10.1) (i.e., of the linearized continuity equa tion), while the addition i7okzVz appears on the lefthand side of equation (10.2) for the ions (i.e, the linearized ion heatbalance equa tion). In addition, we have the inertial equation (9.58). Linearizing
168
Electromagnetic Instabilities in an Inhomogeneous Plasma
the last term in the brackets of (9.58), we find in the smallscale perturbation approximation:
cuihiMiTiodVi/dt + V •tt) = B 0B xk l k z/4irky.
(10.42)
The lefthand side of this equality is calculated precisely in the same manner as in section 10.1. As a result, the inertial equation (10.42) reduces to the equation of form (10.7) with the additional term
 i k zc2A(Kp/u pi)B x/ B 0 in the righthand side. We supplement the abovementioned equations by the pressure balance equation (10.3) which is unchanged. It is obtained, e.g., from Maxwell’s equation for the field B z,
d B z/d y « Airjy/c
(10.43)
and by using the expression for the current density (9.34). In addition, we also.have the electron and ion longitudinal motion equations (9.20) and the second equation of (10.31) relating E z and B x. One can consider R e = 0 in the electron longitudinal motion equation; this is a consequence of the approximate equalities (10.39) and (10.41). Therefore, using the second equation of (10.31), from the electron longitudinal motion equation we find (cf. (10.35)) ("A + iA ^ y V Bo n0 u ; J
(10.44)
By means of this relation one can exclude B x from all the remaining equations. In the ion longitudinal motion equation we allow for the oblique (magnetic) viscosity . Then, using the electron longitudinal motion equation we have the following starting equation
MifidVzi/dt =  d ( p e + p i)/d z  dnza/d x a  ein0(yf*  Vpe)B x/c (10.45) where V*^ = cToiHp/e\Bo is the ion pressuregradient drift velocity. Here it is assumed that wzz * 0. According to (9.30), to the required accuracy the components nxz and iryz have the form _
_ 
_
Pi \dV dz " a
_Pi
9K dz
8Vz 2 ( d q y dqz 5pi V\ fdez + d y j ~dy + 5^ W 2 /d f c dq, dx bp, V dz dx
(1 0 .4 6 )
Equations for Basic Types of Hydrodynamic Perturbations
169
In the righthand side of these equations we have omitted the ion subscripts for V and q. It is assumed that Vj_ and q± are of the form of (9.32) and (9.33). Using equations (10.44) and (10.46), we reduce the linearized equation (10.45) to the form
Now we make the transition to the variables X , Y, Z and B z. Then, as before, we find that B z is contained only in the inertial equation, whilst in all the remaining equations only X , Y, Z appear. Therefore, considering (X ,Y ,Z ) = 0, from the inertial equation we obtain the dispersion relation for the Alfven perturbations which coincides with (4.10). On the other hand, for the perturbations with ( X ,Y ,Z ) ^ 0 the following dispersion relation is obtained [2.2]
+^
+ I) (s'3+ +^ I) (10.48)
Here x = w/w*;, y = (1 + r))~\ and a = k 2 zT0/MiU*?. We shall examine equation (10.48) in section 11.3.
10.3.3 Perturbations in a high/3 plasma with kz > u)/vTl Let us augment the equations of section 10.3.2 by the longitudinal ion viscosity, ttzz ^ 0, and by the longitudinal ion heat flux described by the second term of the righthand side of the first equation of (9.29). When allowing for irzz it is more convenient to use, instead of the perturbed ion pressure p ,, the quantity pi± given by the first equality of (10.15). In addition, when w2Z ^ 0, it is necessary to take into account the modification of the perturbed pressurebalance equa tion described by (10.14). In calculating irzona = ( x ,y ,z ), we use equations (9.30) taking into account the contribution of both the longitudinal heat flux and the part of the vector q* proportional to 1/i/i (see the second equation of (9.29)) in these viscosity tensor com ponents, approximating 25/24 to unity for simplicity. The remaining assumptions are the same as in section 10.3.2.
170
Electromagnetic Instabilities in an Inhomogeneous Plasma
We write the linearized equations in terms of the variables X ±,Y ± ,Zj_ defined by (10.16). As a result, we arrive at the following equation set [10.2]: the electron continuity equation + ( kb
ttn)^JL ~ ojZx + kzVz = 0
(10.49)
the electron longitudinal motion equation (u  u ne)X ± + kt Y± + (u j  L>*pe)Z± = 0
(10.50)
the ion heatbalance equation (a;  27 0u;Di  2iA 1)X ± + ( kp  7oKB)1x + To^Di^i  7fokzVz = 0 (10.51)
the ion longitudinal motion equation (w  u m )kzVz =
+ 2 ^ ® (Xj. + z ±)  i U ^ n X j . k t Mi 5 (10.52) the linearized form o f the first equation o f (9.30) fo r the compo nents o f the viscosity tensor 2M ,n 0
xzz = —i0.96—  kzVz  ( ^ w Di + 2 .9 i A !) x x  ^ k b Yj . + (10.53) and, finally,
the inertial equation (Q i0)  * M )
+ o,Di kzVz  ( «  3u;Di + i^ A x  ^ ) * x
kb ( 3 \ f 3 k 2zc\\ 1 I U  WDi ) Yl — ( rW oi )Z± = 0 . wDi V 2 J \2 ujTi J _
(10.54) Here Ai = (2/3)3.9fc2T0/Mii/i, the quantity was introduced in section 10.2.2. The equation set (10.49)(10.54) will be used in section 12.3.3.
10.3.4 Perturbations in a strongly collisional plasma In this section we shall obtain the dispersion relation for the iner tialess perturbations allowing for the heat transfer, the finiteness of the longitudinal electron heat conductivity and the ion longitudinal
171
Equations for Basic Types of Hydrodynamic Perturbations
motion . Such a dispersion relation will contain the results of sec tions 10.2.3, 10.3.1 and 10.3.2 as particular cases. In addition, it will be suitable for examining oblique perturbations in the strongly collisional plasma regime . We start with the following linearized hydrodynamic equations [10.3]: the electron continuity equation W—  WDe— Tlo
Poe
V, +
 k b ) + (u 
0
B
B
0
= 0 (10.55)
the electron heatbalance equation 7X (~ 7 c M )e +
Pq
 [  27oWDe + i(^.e + A ) ] —
n0
+
lo k z V z 
Poe
cE i^ (« p Bo

Pi
+
Poi
7o«b)
— = 0 + 7o(a, _ 2w; e + u;„e) El _i aA— ^
Bq
kz Bo
(10.56)
the ion heatbalance equation TI
Pq
n0
Poe
7o^Di it'ie— + (w 
+ 1^§l (k p D0
2 7 0u;d,
Pi
+ wie) —  7okzVz Poi
7o«b)  7o(w  2Wpi + Wni)^ = 0 DO
(10.57)
the electron longitudinal motion equation k L l V no
(iM
)^ i^ ^ ) Wpe 5o y
Poe
 i[w + swne  (1 1
^
u^e oo
( 10.5 8 )
= 0
and the ion longitudinal motion equation written with an al lowance fo r the electron longitudinal motion equation My, Here it is assumed that T0j = T0e = T so that p0e = poi, u>*e = —w*,, u>ne = —Uni i WDe = —^Di The quantity A was defined in section 10.3.1.
172
Electromagnetic Instabilities in an Inhomogeneous Plasma
We express Vz and B x in terms of the remaining variables by means of equations (10.58) and (10.59). Then, as usual, we introduce the variables X ,Y ,Z defined by equations (10.4). As a result, we obtain three equations for these variables. Equating the determinant of these equations to zero, we find the following dispersion relation [10.3]: (w  WDi)Q(o;)  k 2 zTR (u)/M i = 0. (10.60) Here « (* )= {(;
■ 5 * V J ■  * ’ ••• 12
+ 2il/ie( 1 + ~P JwJ [ w 
♦!
o(wni + 2u>*j)a>nj +
R( uj) = yO>2 + 2s(u*pi 
r8 w 
+ (1 + s)u;*i]
( i ' 'H )
(u> —
(10.61)
 2vieA + iA
i
*  5w ni\I   0 u pi < + 2ii/ie If wpi VieAM] k lT
u> + 2s \ X
T Wni
+ ^ 0 2Di)
1+
/A
p .
(10.62)
When kz = 0, from (10.60) we obtain (10.26) as well as (10.33) with k\ —►0 and the equation for the transverse ion magnetodrift
oscillation branch uj = c^D j.
(10.63)
When kz ^ 0 but k^v^Ju2 —►0 and the heat transfer is neglected, vie —►0, then from equation (10.60) we find (10.38) and (10.63). When A —> oo and vle —> 0, (10.60) reduces to (10.48) correspond ing to the perturbations with kz ~ uj/vji in the weakly collisional plasma regime .
173
Equations for Basic Types of Hydrodynamic Perturbations
Finally, when one takes vie —►oo, A —> 0 in (10.60), we arrive at the following dispersion relation for the strongly collisional plasma w(w  wDi)[u; + su>ne  (1 + s)u* }
kzVTe 1 + 5/3/6
= 0.
(10.64)
This equation has been found in [10.4] for VT = 0 and in [10.3] for arbitrary V T. We shall use it in sections 11.4 and 12.4.
10.4
Dispersion relations for inertialess perturbations in a plasma flow with an inhomogeneous velocity profile
10.41 Starting equations As before, we describe each particle species (ions and electrons) by their density n, velocity V and pressure p. In addition, the viscosity tensor 7r and the heat flux q are taken into account for the ions. The densities and the velocities of the particles are determined by the equations of continuity and motion (9.19) and (9.20) with R = 0. We assume the electron heat conductivity to be infinite so that the electron heatbalance equation has the form of (10.39). The ion heatbalance equation is written in the form (cf. ( 9 .2 1 )) ^
~ loTi
+ (7o ~ 1^div 9,1 + 7odiv 9 1 =
0
(1065)
where the ion adiabatic exponent 70 is unspecified as yet. For such a description, the vector q± means (cf. (9.33)) x VTl
(1066)
whilst q\\ = e 0q\\ is the longitudinal heat flux which in the case of a collisionless plasma is defined by the equation
( 1 0 6 7 )
This equation can be constructed using the first equation of (9.26) augmented by the term with d q/d t. We use (10.67) when the gradient of the equilibrium pressure is nonessential. Otherwise, (10.67) should be replaced by a more com plex relation.
174
Electromagnetic Instabilities in an Inhomogeneous Plasma
In section 9.3 we have given the expressions for the viscosity ten sor components 7ra/3 in the reference frame for which one of the axes is directed along the magnetic field (see (9.30)). However, for sub sequent calculations we require t t in an arbitrary reference frame. According to [9.3, 9.4], neglecting the dissipative terms, the required form of 7ra/3 is Vap = l r (W3a/3+ 2W ia?)
(1068)
where (10 69) ^ 4a (3 =
f hph^C^^hyW^j) .
Here is defined by (9.25), 6%p = ~ ha hp, and h = e 0 is, as above, the unit vector along the magnetic field.
10.42 Inertialess perturbations in a plasma flow with V n0 =.V T 0 = 0 In this section we assume V n0 = VT0 = 0. Therefore, according to (1.42), V B 0 = 0. With these assumptions the equilibrium velocity of the plasma components (ions and electrons) across the magnetic field vanishes, so Vo = e zU. The velocity U is assumed to be dependent upon x ,dU /dx 0. It follows from (10.68) that at dU/dx ^ 0 the equilibrium values of the viscositytensor components, iryz and 7r*y, differ from zero _ Poi dU
TTozy  *0yz 
dx .
/m (10.70)
Linearizing the equations of motion (9.20), for each particle species we obtain (cf. (9.32)) v,
= L (
e
, + V
Bo \ *
c
b
,  ’M
)
enQJ
(10.71)
Vy = ^  (  E x +  B y + — ^ Bo \ c en0 ox When VT0 = 0, linearization of (10.66) yields simply
q± = 0.
(10.72)
Allowing for (10.71) and (10.72), the linearized ion equations of continuity (9.19) and heat balance (10.65) take the form (cf. (10.1),
Equations for Basic Types of Hydrodynamic Perturbations
, n
 u — f
n0
u (^ ~ \Poi
Bz Bo
iB x dU 5 Bo dx
7 o— ^
+ k z ( 70 
uj— ^   — +
n 0J

kzVz =
^
1)—
poi
=
175
0
( 1 0 .7 3 )
0
where uj = uj — kzU. According to (9.20) and (10.67), Vz and qz satisfy the equations
\n0 (  iuj V z
I
^ “j ~ )
• V • 7T =
+
\kzpx ( en0E z (10.74)
 i ujqz = — —— ( Po\^z •V •7r  ifcy7r02j,pi J. Mi»0 V / As follows from (10.68), the perturbed viscositytensor compo nents contained in these equations are equal to (note— only the es sential terms are written) 
y
ipoi
_
ZX —
f i
(10.75)
Poi ^ L + ^ f A _ ^ i V
dx
dx \poi
B 0)
and f zz = 0, while ir0zy is determined by (10.70). Using these ex pressions, we obtain the relations T. d U , in° ' d i + _
€z V
.
• 7T
Min od U f _ , JTif ' v •71= \ ’ + •_ _
ifcy 7To*y
.
Pi
—
Toi f.\
ikypoiB z dU O D
^1
(10.76) Note that in the first equation of (10.76) the terms with pidU /dx, contained in the product VxdU /dx and in the divergence of the vis cosity tensor, are mutually compensated . Allowing for (10.76), it follows from (10.74) that
176
Electromagnetic Instabilities in an Inhomogeneous Plasma
kzQi dx *
(1 0 .7 8 )
Substitution of (10.77) into (10.73) yields
(10.79) Note that the term with B z in the second equation of (10.79) is due to the longitudinal heat flux qz. Neglecting qz, from the second equation of (10.79) the trivial result pi = 7 oTo^ should follow. We supplement (10.79) by the pressurebalance equation (10.3) and by the relations
Te = 0 ^kzT()en J eeU()Ez — 0
(10.80)
following from (10.39) and (9.20). For the case where CS1 > A^T0i/Mi, the dispersion relation follows from the above equations:
(10.81)
Here, as usual, /?i>e = 87rn0 Toi)e/ B$. We shall use equation (10.81) in section 13.5.1.
10.43 Allowance fo r the density gradient Let us supplement the results of section 10.4.2 by the terms with Vn0. For finite /?, according to (1.42), it is also necessary to take into account VjE?0 in addition to Vn0. Describing the ions, we shall follow the isothermal approximation, 70 = 1 , neglecting the ion tem perature perturbations, TJ = 0 , and the ion heat flux, q x = 0 . By analogy with (10.73), we reduce the continuity equation to the form
 ( u  u ) Di)n + ( t i  u ni) B  i e E z(u>ni v Di) /k zT0i + kzVz = 0 . (10.82)
Equations for Basic Types of Hydrodynamic Perturbations
177
Here new variables are introduced: n  — _ 1Kn Bx
Hq
B =
kz B q
ikb B x Bo
(10.83)
kz B 0
V, = Vz  L ^ L h . kz dx B 0' The ion longitudinal motion equation now yields /(—
ntV ie r, /, \ k2 zTo\ . OtlkzTo\ a c D,)K = H £ , ( l  oO + i j  n  —
(10.84)
We supplement (10.82) and (10.84) by the equations
B + fih / 2 = 0 e eE z  ikzT0eh = 0
(10.85)
which are similar to (10.3) and (10.80). By means of (10.82)(10.85) we obtain the dispersion relation
(i +f) ("“*>(*ft:"" k 2T(Oe Mi
i
,
2a
/
/? r oi \
( 10.86 ) = 0.
We shall consider equation (10.86) in section 13.5.3.
10.5
Description of Alfven perturbations in a plasma flow with an inhomogeneous velocity profile
To obtain equations for the Alfven perturbations in a plasma flow with an inhomogeneous velocity profile it is necessary to add one more independent equation to the equations of section 10.4. We take the inertial equation (9.58) as such an equation.
10.5.1 Alfven perturbations when neglecting drift effects Let us neglect the viscosity in (9.58). Calculating the perturbed velocities (see (10.71)), we neglect the ion perturbed pressure so that V± = c E x e z/ B 0 + U B ± /B 0. (10.87)
178
Electromagnetic Instabilities in an Inhomogeneous Plasma
We neglect the last term in the derivative d/d t = d /d t + TJd/dz + V± V as being of the order of u* / uj. In addition, we neglect the perturbation B z so that B = [Bx,( i/k y)dB x/dx,0]. The equilibrium magnetic field is assumed to be uniform, VJ30 = 0. Finally, assuming E z = 0, we express E± in terms of B x
E = i u dfir ckykz dx
(10.88)
E = As a result of this, we obtain that (9.58) reduces to V i •[(£2  c2 Ak 2z)V ±B x) = 0.
(10.89)
This is the starting equation for the ordinary KelvinHelmholtz instability. We shall examine (10.89) in section 13.1.
10.5.2 Magnetic viscosity tensor fo r dU/dx fi 0? with B x ^ 0 To add the drift effects to (10.89) it is necessary to have expressions for 7rxx,7Tyy,7rxy = ttyx^ zx^yz* The two last quantities are given by (10.75) and (10.70). We find the remaining components of the viscosity tensor by linearizing (10.68) with respect to hx and h y: ^xx 7Tyy =
»y
P (0 V X 2fi \ dy
OVy dx
BydU B 0 dx
n B 0 dx p ( dVx , B x dU ~ 2 ( l\ d x + Bo dx
(10.90)
xx^
8% dy
Note that to obtain B ± dU /dx it is necessary to use (10.68) and (10.69) but not (9.57). In [10.8], equation (9.57) has been used in correctly.
10.5.3 Allowance fo r the drift effects fo r Vn0 = VT0 = 0 Let us reduce (9.58) using the fact that the variables n, B , and Vz defined by (10.83) vanish in the case of Alfven perturbations. As in section 10.4.3, we assume VT0 = 0. In addition, we take Vno = 0. Then, according to (10.83)
Equations for Basic Types of Hydrodynamic Perturbations
179
In this case p and fi in (10.90) should be considered to be equal to their equilibrium values independent of the coordinates. The terms with pi and B z should now be omitted in (10.75). In (10.71), as in section 10.5.1, we omit the term with p and express E x , E y, and By in terms of B x by means of equations (10.88) and of the relation B y = (i/ k y)d B x/d x . Then, as in section 10.5.1, we have
K = v = y~
h' B° a i 1U> x kykzB 0 dx '
n°.9'2)
Using (10.75) and (10.90), we find curl* V •7r =
72rv T> , 1 d (? , d2U V j V •Vj. + — ^ \ B X
B odx\
, i 2 B x dU yB 0 dx.
ikzPoi
dx2
d B x dU dx dx
,
(10.93) Further transformation of (9.58) using the above relations leads to V x •[(u2  c\k2 + a\k2To\lM\)S7±_BX\= 0
(10.94)
where a\ is defined by (10.78). Equation (10.94) will be used in the analysis of section 13.3.
10.5.4 Dispersion relation o f smallscale Alfven perturbations fo r Vho = VTo = 0 When \d\nBx/dx\ > dln?7/dx, the dispersion relation follows from (10.94): u 2  c\k2z + a\k2 zT0\IM\ = 0. (10.95) Exactly the same dispersion relation is also obtained in the kinetic approach (see (2.5), (2.7) and the first equation of (2.26)). The consequences following from (10.95) will be considered in sec tion 13.4.1.
10.5.5 Dispersion relation o f smallscale Alfven perturbations fo r V n 0 7^ 0, with VT 0 = 0 Restricting ourselves to the case of smallscale perturbations, we re duce (9.58) to (10.42). In contrast to section 10.5.4, we now consider Vn 0 ^ 0 and V B 0 ^ 0. By analogy with section 10.5.3, we find that
180
Electromagnetic Instabilities in an Inhomogeneous Plasma
the righthand side of (10.93) is augmented by the term £{curl* V7r}, where ^ c u r l ,V ..} = A (  ) v i V ', .
(10.96)
The procedure for obtaining the dispersion relation by means of (10.42) reduces to expressing all the perturbed quantities in terms of B x making allowance for the quantities (10.83) which vanish, i.e., (cf. (10.10), (10.91)) 71 lKfi B x
n0
k z Bo
Ih = WflS* B0 K B0 V =
(10.97)
kz dx Bo ’
Then the part of curl^V •7r given by the righthand side of (10.93) is an unchanged function of B x when Vn0 ^ 0, VJ50 ^ 0, which can be proved by means of the expression for the perturbed velocity
 v
( i o
9 8 )
and by equations (10.97). (Here, as above, V„ = cT0Kn/ e B 0.) The righthand side of (10.96) is expressed in terms of Bx using (10.97), (10.98) and (10.88). Then
Vx = (£>  u ni)B x/ k zB 0.
(10.99)
Finally, the quantity Vy contained in (10.42) under the operator d/d t = —i(w —u>ni) can be replaced by (i /k y)dVx/d x with the help of the approximate condition div V = 0. As a result, from (10.42) we find the dispersion relation (w  u;ni)(u>  u;Dj)  c\k\ + a^ T oi/M i = 0.
(10.100)
This equation coincides with that obtained by the kinetic method (cf. (2.5), (2.7), (2.9), and (2.26)). We shall use (10.100) in section 13.4.2.
11 Instabilities in a Collisional Finite/? Plasma
Using the results of Chapters 9 and 10 as starting equations, we shall now analyse the lowfrequency longwavelength instabilities in a collisional finite/? plasma. This analysis will also be continued in Chapter 12 where the material related to a plasma with /? > 1 and d In p /d In n /vTi, i.e., to those for which the ion longitudinal motion is important. The plasma is considered to be weakly collisional in section 11.3 and strongly collisional in section 11.4, in the meaning given in Chapters 9 and 10. The Alfven (inertial) instabilities in the collisional regime are discussed in sec tion 11.5. There we deal with only the dissipative effects of order k 2L p2u J u since, as mentioned in Chapter 10, the terms describing the remaining dissipative effects are not contained in the dispersion relations for these instabilities. Regarding the classification of the types of perturbations dis cussed, sections 11.111.4 are similar to Chapter 3 where the iner tialess perturbations in a collisionless plasma have been considered, 181
182
Electromagnetic Instabilities in an Inhomogeneous Plasma
whereas section 11.5, similar to Chapter 4, is devoted to inertial perturbations in such a plasma. Comparing the collisionless and collisional regimes, one can find what properties the perturbations in these regimes have in common. It is clear that the collisionless regime should first of all be compared with the weakly collisional regime. This is just the aspect which we shall consider below. It is also of specific interest to compare the results of the weakly and strongly collisional regime as discussed below. According to section 11.1.1, in a weakly collisional plasma there are two branches of the inertialess magnetodrift entropy oscillations which grow when the temperature gradient is compared with the density gradient. In this respect we have an obvious analogy with Tserkovnikov’s magnetodrift instabilities in a collisionless plasma discussed in section 3.1. As for the dissipative effects of order u/u\ and (k^piYui/tJ discussed in sections 11.1.2 and 11.1.3, they cause the entropywave instability even for VT = 0. This is evidence of the fact that a collisional plasma is more unstable than a collisionless one. At the same time, according to section 11.1.4, the entropy waves, as well as Tserkovnikov’s waves, exist only for not too a large collision frequency, vie < u ( the weakly collisional regime). In the opposite case, when uie > u;, we have two essentially dissipative oscillation branches instead of the entropy waves. One of those branches is ex cited under the same temperature gradients as the entropy waves, but its growth rate becomes small as the ratio u /v le. Note that such a sensitivity of the entropy waves to the degree of collisionality is explained by the fact that, when vlQ> u , the collisional heat trans fer prevents differences in the perturbed temperatures of different particle species as required for these waves. Recall that in section 3.2 we examined the transverse drift waves in a collisionless plasma. However, as follows from section 10.1.4., this oscillation branch exists also in a collisional plasma. In section 3.2 we allowed for the small terms of order M e jM\ and observed the transverse driftwave instability with the growth rate of the abovementioned order for VT = 0. That instability has been of interest since all the remaining oscillation branches become stable for VT = 0. However, as is clear from the above, in a collisional plasma there are oscillation branches which are unstable due to dissipative effects even for VT = 0. Therefore, in such a plasma the effects of order M e/M [ are not of interest, and we do not consider them. As mentioned in section 3.3, when k2 increases to magnitudes of order u /v Te and higher, in a collisionless plasma , first, the iner tialess magnetodrift branches become coupled with the transverse drift waves, and, second, one of the magnetodrift branches disap
Instabilities in a Collisional Finiteft Plasma
183
pears due to the broadening of the electron magnetodrift resonance by the electron longitudinal thermal motion. As a result, in that case we had two branches of inertialess perturbations: the electron drift branch and the single magnetodrift branch. Those branches can be excited only in the presence of a temperature gradient. The analysis of section 11.2 shows that similar modification of the inertial pertur bation branches also holds in a collisional plasma . However, some differences are now revealed. One of them concerns the characteristic longitudinal wave number for which the above modification occurs. This wave number becomes of order kz ~ (uji/ q)1^2/v Te which is es sentially larger than that in the case of a collisionless plasma. The given kz increases with increasing collision frequency. When vie ~ i.e., just as under the condition when the heat transfer is important (see above), this kz becomes of order u / vt\. In this case the ion longitudinal motion is important. At the same time, the condition vie ~ u corresponds to the strongly collisional plasma regime . There fore, the abovementioned analogy in the behaviour of the oscillation branches concerns only the weakly collisional plasma regime. Another essential difference in the behaviour of the inertialess branches with kz ^ 0 in the collisional and collisionless regimes is as sociated with difference of the temperature gradient ranges for which these branches are unstable. In particular, the perturbations with such kz in the collisional plasma become unstable even for VT = 0. Such an instability is due to the finite electron longitudinal heat con ductivity. Therefore, although the effect of the electron longitudinal heat conductivity is similar to that of the electron longitudinal ther mal motion, there is no complete analogy here. The analogy between properties of the inertialess perturbations in the collisionless and collisional plasma is reestablished in the limit of infinitely large electron heat conductivity. This approximation can be used, in particular, in the case where kz c± u / vt\discussed in section 11.3. It then appears that, if VT = 0, both a weakly collisional plasma and a collisionless one are stable against the perturbations considered. Summarizing, we conclude that inertialess perturbations in a weakly collisional plasma (when neglecting dissipative effects) are very similar to those in a collisionless plasma. Besides the rest, this analogy is also of methodological interest since, if necessary, a more complicated kinetic description can be replaced by a hydrodynamic one without an essential loss of information. However, there is no complete analogy between dissipative effects in a collisional plasma and kinetic resonant effects (collisionless dissipation) in a collision less one. The inertialess perturbations in a strongly collisional plasma have
184
Electromagnetic Instabilities in an Inhomogeneous Plasma
essentially different properties. According to section 11.4, they are excited for all (3 even in the case where VT = 0. This can be considered as one more evidence of the fact that the collisions as a whole favour a destabilization of the inertialess perturbations. The role of collisions is more passive in the inertial perturbations. According to section 11.5, when VT = 0, the dissipative effects do not result in their excitation for all /3. It should also be remembered that, as shown in Chapter 4, when VT = 0, the inertial perturbations are stable also in the absence of collisions. The above remarks do not concern forcefree or almost forcefree high/? plasmas which will be discussed in Chapter 12. The results presented in this chapter have been obtained in [2.1, 2.2, 3.1, 9.1, 9.9, 10.310.5, 11.111.3]. The specific contributions of these papers will be mentioned when considering the respective problems. Nonlinear effects related to the corresponding types of perturba tions were considered in [10.7, 11.411.6]. As a rule, in the present chapter and in Chapters 12 and 13 we omit the subscript zero in the equilibrium number density and mag netic field, i.e., n0 —►n,2?0 B.
11.1
Magnetodrift entropy instabilities
11.1.1 Entropy perturbations when neglecting dissipative effects According to section 10.1.1, the entropy perturbations when neglect ing dissipative effects are described by dispersion relation (10.6). From (10.6), for Te = T\ = T the instability criterion follows [3.1, 2 .1]
(11.1) In particular, when (3 oi 1, (11.11) has the solution
u = 2wiD i'
1 + 5/3/12 1 + 2/3/3 ’
( 11.12)
This corresponds to the magnetodrift branch similar to that dis cussed in section 3.3.2.
Instabilities in a Collisional Finite/3 Plasma
187
As a collisional analogue of (3.28) describing the connection be tween the electron drift and magnetodrift branches for VT ^ 0, the following dispersion relation is obtained from (10.38) for A —►oo: w2 ( 1 + ^7o/3 )  u
Wne i 1 + ^2~) +
+ 7ouwu>Dl( ^
\ 2 Kp
+
1
 ^
Kn

7oU,D ii 1 + 2
5^
 )
270 k b )
=
(11.13)
0.
It follows hence that (3.29) is valid also for VT ^ 0 if (3 wDj, it follows from (11.15) that unstable perturbations have the frequency and the growth rate defined by (cf.
(8. 12)) Rew = 7
= u>Di/9A .
(1 1 .1 7 )
188
Electromagnetic Instabilities in an Inhomogeneous Plasma
The instability described by (11.11) does not hold for too large (3. Equating the real and imaginary parts of ( 1 1 . 1 1 ) to zero at 7 = 0, we find the instability boundary A) =
2/ 70
= 6/5.
(11.18)
When (3 > /30, the plasma is stable. It follows from the foregoing analysis that a plasma with VT = 0 and finite f3 is unstable. The instability is due to the fact that in plasma with finite, albeit small, f3 there is a branch of sufficiently slow oscillations, u / k y ~ Vb, i.e., the magnetodrift oscillations , which are excited as a result of the negative dissipation related to the finite electron heat conductivity . The longitudinal wave number of this instability is of order
kz ~ (^eWne/?)1/2/vTe
(11.19)
The maximum growth rate is numerically small compared with the frequency 7
~ O.IReu; ~ 0.1fi(ckyT /eB )d \ n n /d x .
(11.20)
Note that the nature of the excitation is essentially associated with the temperature perturbations , whereas the presence of an equi librium temperature gradient is not required.
11.2.2 Instability o f electron drift waves in a plasma with VT / 0 By analogy with section 3.3.1, we now consider the influence of a small temperature gradient on the branch (3.29). Assuming rj < 1 and keeping the small terms of order rj in (10.38), we find the growth rate of this branch 7
7
=
WA ^ 1
+
+ ^
+ 7°^ + 7o 0, ( 1 1 .2 1 ) describes the instability due to the finite electron heat conductivity initially examined in [3.23] (see also [1 1 .7 ]). This instability was discussed in [2], section 5.3.
Instabilities in a Collisional Finite,] Plasma
11.3
189
Inertialess instabilities with kz ~ ^'/vTl in a weakly collisional plasma
Let us consider the instabilities described by (10.48). At first, fol lowing [2 .2 ], we shall present an analytical examination of equation (10.48) and then the results of the numerical calculations of [2 .2 ] refined in [11.1, 11.3] will be given. When /? —*•0, (10.48) means [11.8, 11.9]
x
_3
(& 2 /3  rj 0 * + ^'
z 3 + — a
I + 77
\3
1+ v J
=0.
(11.22)
The consequences of (11.22) were discussed in [2], section 5.3. This equation describes unstable perturbations when tj
> 2 /3
or
77
< 2 .
(11.23)
For /3 > 1 and 1 + r? > 1//3, in the leading order in 1//3, all the roots of (10.48) are real. Two roots coincide for a = (5/4)(/3/4)3. In the next order in 1//3, we find that the coinciding roots are complex when t? > 1. (11.24) Then
5
Reu> =  ^ w Di R ew 
/ j )  l \ 1/2
(1 L 2 5 )
7 _ 2(3/?)i/2 U + i ; ' Thus, it can be seen that a plasma with a sufficiently large pos itive temperature gradient is unstable for both small and large (3. However, increasing f3 from a small values up to large values, the instability range of a plasma with 77 > 0 is not continues. This range has a gap for /3 ~ 1 (see figure 11.1), [2 .2 ]. The instability region for a plasma with 77 < 0 is also nontrivial (see figure 11.1). It can be seen that the instability described by the second inequality of (11.23) disappears for finite f3 < 1. Instead, for j3 = 1.2 a new instability appears. Its boundary is displaced, with increasing (3, from 77 = 0 (for (3 = 1 .2 ) up to 77 = —1 (for f3 = 0 0 ), [2 .2 ], (see figure 1 1 . 1 ). In addition, as seen from figure 1 1 . 1 , there is one more stability gap in the region of negative 77 [11.1, 11.3]. As for a plasma with VT = 0 , it is stable against the discussed perturbations for all (3.
190
Electromagnetic Instabilities in an Inhomogeneous Plasma
F igu re plasma.
11.4
11.1
Instability
ranges
(shaded)
of collisional
Inertialess instabilities in a strongly collisional plasma
The inertialess perturbations in a strongly collisional plasma are described by dispersion relation (10.64). We restrict ourselves to analysing it for VT = 0 [10.3, 10.4]. An analysis of (10.64) for arbitrary V T was carried out in [10.3, 11.2]. When V T = 0, (10.64) takes the form x ( x + 0 ( x + l )  4 a ( a; y ) / ( ^ +  ) = 0
( 1L26)
Here x = u /u nna = k 2 zT / Thi s equation describes the socalled thermalforce instability which occurs for both small and large (3. For small (3 this instability was initially studied in [11.10] and then in [11.8, 11.9]. The corresponding results of these papers are presented in [2], section 5.4. When (3 > 1, one of the roots of (11.26) is large (it corresponds to stable oscillations) while the two other roots satisfy the equation * (* + l )  ^ ( *  y ) = °
( 1L27)
Hence it follows that there is an instability if 0.2 < a 1/2//? < 0.6.
(11.28)
Instabilities in a Collisional Finite/? Plasma
191
This means that the characteristic kz for the unstable perturbations increases with increasing (3,kz ~ /?. The oscillation frequency and the growth rate of the thermalforce instability are of order u n[.
11.5
Inertial instabilities
In analysing the inertial instabilities, we restrict ourselves to the case of perturbations with kz = 0. According to section 10.1.2, in neglecting dissipative effects, the dispersion relation of the inertial perturbations with kz — 0 reduces to (4.1) as has been examined in section 4.1.1. Following from section 10.2.1, dissipative effects of order do not make a contribution in the dispersion relation for such perturbations. The same concerns the heat transfer (see section 10.2.3). Therefore, we should consider only the effects of order (k^piYuJuj. For this, we use dispersion relation (10.22). We analyse it in the case where /? ;)
(11.30) J'i
This means that the perturbations corresponding to the branch neither grow nor are damped while those corresponding to the u 2
branch are damped.
12
Instabilities in Forcefree and Almost Forcefree High/? Plasmas
The study of instabilities in inhomogeneous forcefree or almost forcefree high(3 plasmas (i.e., in those with vanishing or almost vanishing pressure gradients, Vp « 0, and with /? > 1) is of interest, in particular, in connection with the problem of hotplasma confine ment due to the pressure of a dense blanket o f neutral gas . This problem was first discussed in [12.1] and then in [12.212.5] and in a number of other papers. As noted in [12.6], similar conditions can be realized in a laserproduced plasm a . In [2.6, 12.6] it has been men tioned that the case where /? >> 1 and Vp ~ 0 is typical for many space situations, in particular, for certain solar wind regions. Interest in forcefree high/? plasma has also increased in connec tion with the problem of plasma compression by imploding liner (see the literature on this problem, e.g., in [12.7, 12.8]). It is clear that plasma with /? > 1 and a small pressure gradient can be realized in the central region of thetapinches , in usual mirror machines as well as in the tandem mirror experiments ( t m x ) [6.14, 6.15], and in the reversed field configurations (see, e.g., [12.9, 12.10]). The above situ ation is typical also for the problem of magnetically insulated inertial confinement fusion discussed in [12.1112.13]. Using the basic equations concerning forcefree or almost forcefree high,/? plasmas given in Chapters 2 and 10, we shall examine inertialess instabilities in such plasmas in sections 12.112.4 and the inertial (Alfven) instabilities in section 12.5. In sections 12.1 and 12.2 the plasma is assumed to be collisionless . Then in section 12.1 it is considered to be strictly forcefree, V B — 0, while in section 12.2 it is almost forcefree , Vp « 0, V B ^ 0. In con trast to sections 12.1 and 12.2, in sections 12.3 and 12.4 a collisional 192
Instabilities in Forcefree and A lm ost Forcefree High/3 Plasmas
193
plasma is studied. We deal with a weakly collisional plasma in sec tion 12.3 and with a strongly collisional one in section 12.4 (within
the meaning explained in sections 10.3.2 and 10.3.4) In section 12.5, in analysing the Alfven instabilities we allow for both collisionless and collisional effects. When the problem of plasma confinement by a neutral gas blanket is kept in mind, in studying instabilities it is necessary to distinguish the central region , where the plasma has maximum temperature and minimum density, and the peripheral region, where the temperature is low and the density is high. Recall that the characteristic fre quency of gradient (drift) perturbations increases with temperature (u;* ~ T) and the binary collision frequency increases with density, but decreases with increasing temperature. Then it is clear that the theory of gradient instabilities in the peripheral region must clearly allow for collisions while the central region can be described in some cases by the collisionless approximation. In accordance with Chapters 1 and 2 (see sections 1.2.3 and 2.1.4), the forcefree plasma (V p = V B = 0 ) is of interest from the physi cal viewpoint since in such a plasma there are no particle magnetic drift and Larmor (diamagnetic) current. That is why at first sight it can seem that there is no reason for instabilities. However, on more careful consideration of the equilibria of forcefree plasma it can be seen that, due to the presence of temperature and density gradients (when V p = 0 , these gradients are related by the expres sion V(nT) = 0) in such a plasma, there are higherorder fluxes , e.g., heat flux . It is these higherorder fluxes which are the reason for instabilities of forcefree plasma. Hence it follows, in particu lar, that in a hydrodynamic description of instabilities of forcefree (and almost forcefree) plasmas, generally speaking, it is necessary to use the equations correctly allowing for the higherorder moments of the distribution function. Then instabilities in weakly collisional and collisionless forcefree plasmas are similar (cf. section 12.1 with section 12.3, and also section 12.4.1 with section 12.4.3). A common property for both cases is an instability of the oblique ( kz ^ 0 ) in ertialess perturbations for (3 —> oo and Te = 7] and the stabilization of these perturbations for finite 1//3 and Te ^ T[. It is necessary to keep in mind that when (3 > 1, the presence of even a small equilibrium plasma pressure gradient causes a signifi cant equilibrium magneticfield gradient (this follows from the equi librium condition (1.42)). Therefore, instabilities in an almost forcefree plasma must be investigated making allowance for the magneticfield gradient and the particle magnetic drift associated with it. We follow this procedure in section 12.2 and in certain parts of sections 12.312.5. Then the picture of instabilities depends very essentially
194
Electromagnetic Instabilities in an Inhomogeneous Plasma
on the presence of a magneticfield gradient and on the direction of
this gradient. As a whole, we find that collisionless and weakly col lisional plasmas with d In B / d Inn < 0 are more stable than those with magneticfield and density gradients in the same direction. The condition d l n B / d l n n < 0 means, for the problem of hotplasma confinement by the pressure of a dense blanket of neutral gas , that the magnetic field is greater at the centre than at the pe riphery. It is precisely this type of magnetic system which is most appropriate to the case where the field is used to reduce thermal con ductivity and diffusion but not to contain the plasma. In principle, the presence of a magnetic field is essential only at the centre and not at the periphery where, due to frequent collisions, the magnetization of the plasma is of no importance. Hence it is on the whole a fair con clusion that a plasma with d In B / d Inn < 0 has enhanced stability, though we must not forget that even in this most favourable case the plasma is still not entirely free of instabilities, especially, if the role of the collisional dissipative effects is not negligible. As for a strongly collisional plasma, according to section 12.4, it is most unstable just for d In B / d Inn < 0 , although the growth rate of the perturbations is small as (3~1^2 in this case (it is assumed that (3 > 1 ). In contrast to the case where VT = 0 dealt with in Chapter 11 , the Alfven instabilities can play a more active role for the case V p « 0 discussed in the present chapter. According to section 12.5, such instabilities can be excited even if resonant and dissipative effects are neglected, if —4 < d in B / d l n n < 0. The abovementioned effects enhance the instability region (see, for more detail, section 12 .5 ). On the whole, forcefree and almost forcefree plasmas are more unstable than a plasma with VT = 0. However, it should be noted that in the most unstable case the coefficients of heat conductivity and diffusion caused by the instabilities discussed cannot essentially exceed the Bohm values. The results presented in this chapter have been found in [2.6, 3.5, 3.12, 9.9, 10.110.3, 12.14]. In addition, papers [11.111.3] are associated with the topic of this chapter, where the boundaries of instabilities for various values of (3 and rj in various collisionality regimes were calculated. In [12.15] the shortwavelength perturba tions (k±pi > 1 ) were studied. A part of the results of [2.6, 3.12] has been reproduced in [12.6]. Paper [12.6] was critiqued in [12.16]. However, this critique as well as the results of [12.16] are wrong. Instabilities in a collisional high/? plasma with kz — 0 were also studied in [12.17]. However, the author of [12.17] did not consider a difference in the perturbed temperatures of electrons and ions. Therefore, the growth rate of perturbations considered in [12.17]
Instabilities in Forcefree and A lm ost Forcefree High/3 Plasmas
195
wrongly depends on the difference z — 1 where z is the ion charge number. Nonlinear effects related to the above instabilities were discussed in.[12 .8 ]. 12.1
Inertialess instabilities in a forcefree collisionless high fi plasma
The lowfrequency perturbations (u> t \)2 +
= 0
(12.5)
w = 3 ( i ± ^ I 7 j ) w Ti
( 1 2 .6 )
uj2/2
with complex solutions
It is clear that an instability occurs. This is the hydrodynamic insta bility of a forcefree plasma (it is not associated with the resonant particles). It follows from (12.4) that an instability also occurs for x < 1 when we have [3.12] Reu; = u>t J 2 _ 7T 2 7 2 ^ 24  kz  vTi'
(12.7)
This is the kinetic instability of a forcefree plasma. The collisionless instabilities described by (12.6) and (12.7) have collisional analogues which will be considered in section 12.3. 12.1.2 Plasma with Te ^ T\ and /3 —> oo According to section 12 .1 . 1 , the perturbations in a plasma with —> oo and Te = TI with v?\ Di) —k2zc \ = 0
(12.52)
while (12.49) remains valid. Consequently, in the zeroth approxima tion in 1 //?, a strongly collisional plasma with V B / 0 as well as with VJ5 = 0, is stable against inertial perturbations. Let us now allow for the terms of order 1/(3 assuming V B ^ 0 . Such terms are substantial when one of the solutions of (12.52) coincides with (12.49). Such an intersection of the roots takes place if (.kzcA/oj ni)2 = s(s  b). (12.53) Allowing for the terms of order 1//3, we find that the frequencies of intersecting roots are shifted by a magnitude Su, where, according to (10.64), (