Conformal Field Theory.
 9789812831972, 9812831975

Table of contents :
Chapter I Conformal Symmetry and Fields
I.1 Conformal Invariance
I.2 Symmetries and currents
I.3 Operator product expansion
I.4 Central charge and Virasoro algebra
I.5 Free bosons and fermions on a plane
I.6 Conformed ghost systems
Chapter II Representations of the Virasoro Algebra
II. 1 Fields and states
II. 2 Correlation functions
II. 3 Conformal bootstrap
II. 4 Null states
II. 5 Fusion rules
II. 6 Coulomb gas picture
II. 7 Ising and other statistical models
II. 8 Operator algebra and bootstrap in the minimal models
II. 9 Felder's (BRST) approach to minimal models. Chapter III Partition Functions and BosonizationIII. 1 Free fermions on a torus
III. 2 Free bosons on a torus
III. 3 Chiral bosonization
III. 4 Chiral bosonization on Riemann surfaces
III. 5 BRST approach to minimal models on a torus
Chapter IV AKM Algebras and WZNW Theories
IV. 1 AKM algebras and their representations
IV. 2 Sugawara-Sommerfeld construction
IV. 3 WZNW theories
IV. 4 Free field AKM representations
IV. 5 AKM characters, and the A-D-E classification of modular invariant partition functions
Chapter V Superconformal and Super-AKM Symmetries. V.1 Superconformal algebras and their unitary representationsV. 2 Super-AKM algebra and its representations
V.3 Supersymmetric WZNW theories
V.4 Chiral rings and Landau-Ginzburg models
Chapter VI Coset Models
VI. 1 Goddard-Kent-Olive construction
VI. 2 Kazama-Suzuki construction
VI. 3 KS construction and LG models
VI. 4 Gauged WZNW theories
VI. 5 Felder's (BRST) approach to coset models
VI. 6 Generalized affine-Virasoro construction
Chapter VII W Algebras
VII. 1 W3 algebra and its generalizations
VII. 2 Free field approach
VII. 3 Quantum Drinfeld-Sokolov reduction. VII.4 W coset constructionChapter VIII Conformal Field Theory and Strings
VIII.1 Bosonic strings in D d"26
VIII.2 Supersymmetry and picture-changing
VIII.3 Extended fermionic strings
VIII.4 W gravity and W strings
Chapter IX Quantum 2d Gravity, and Topological Field Theories
IX.1 Quantum 2d gravity
IX.2 Liouville theory
IX.3 Topological field theories and strings
Chapter X CFT and Matrix Models
X.1 Why matrix models?
X.2 One-matrix model solution
X.3 Multi-matrix models
X.4 Matrix models and KdV
Chapter XI CFT and Integrable Models
XI.1 KdV-type hierarchies and flows. XI. 2 W-flowsXI. 3 WZNW models and Toda field theories
XI. 4 4d self-duality and 2d integrable models
XI. 5 Self-duality and supersymmetry
Chapter XII Comments
XII. 1 About the Literature
XII. 2 Basic facts about Lie and AKM algebras
XII. 3 Basic facts about V-functions
XII. 4 Basic facts about Riemann surfaces
XII. 5 Basic Facts about BRST and BFV
XII. 6 CFT and quantum groups
Text Abbreviations

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CONFORMAL FIELD THEORY Sergei V. Ketov Institutfur Theoretische Physik Universitat Hannover Germany

World Scientific Singapore • New Jersey • London • Hong Kong

Published by World Scientific Publishing Co. Pte. Ltd. P O Box 128, Fairer Road, Singapore 912805 USA office: Suite IB, 1060 Main Street, River Edge, NJ 07661 UK office: 57 Shelton Street, Covent Garden, London WC2H 9HE

First published 1995 First reprint 1997

CONFORMAL FIELD THEORY Copyright © 1995 by World Scientific Publishing Co. Pte. Ltd. All rights reserved. This book, orparts thereof, may not be reproduced inanyformor by any means, electronic or mechanical, including photocopying, recording or any information storage and retrieval system now known or to be invented, without written permission from the Publisher.

For photocopying of material in this volume, please pay a copying fee through the Copyright Clearance Center, Inc., 222 Rosewood Drive, Danvers, MA 01923, USA. In this case permission to photocopy is not required from the publisher.

ISBN 981-02-1608-4

Printed in Singapore by Uto-Print

To Tatiana Ketova, my wife, for all her love and support

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Preface The aim of physics as a science is to describe the most fundamental laws of Nature. Being the most fundamental, the physical laws are simultaneously the simplest ones. To be simple does not mean however to be trivial. The simplicity of fundamental physics exhibits itself in fundamental symmetry principles. It is therefore fundamental physical symmetry principles that we should be looking for, and it is mainly a theoretical problem. It is very often in physics that underlying symmetries are hidden by the complexity of actual physical phenomena, and it is the objective of theoretical physics to uncover them. Taking high energy physics as an example, the theoretical basis for our current under­ standing of properties and dynamics of elementary particles is provided by quantum field theory (QFT). A general QFT however seems to be too general for these purposes. It is the gauge QFT's that are apparently distinguished in describing fundamental physical interactions of Nature. Compared to a general QFT, quantum gauge theories are much more restrictive and possess additional symmetries. The well-known achievements of gauge theories in high energy physics during last decades perfectly illustrate the power of symmetry principles in physics. The higher an energy scale, the more symmetric become interactions among elementary particles. This is a clear sign of some (normally broken) symmetry which should be formally restored at infinite energies or vanishing masses. Another relevant example is provided by the theory of critical phenomena. It is a well-known fact that quite different physical systems at their criticality exhibit univer­ sal behaviour. This also indicates a presence of some underlying fundamental symmetry. It is the conformal symmetry that is responsible in both cases. One important point should however be emphasized. Benefits of having conformal symmetry usually come true in two dimensions, where this symmetry becomes infinitedimensional. An infinite number of conservation laws is responsible for an integrability of two-dimensional physical systems possessing conformal invariance. It is the twodimensional conformal field theory (CFT) that is actually addressed in this book. CFT is viewed as a specific QFT having conformal symmetry and its own general principles. The underlying theoretical principles of CFT were actually extracted from a careful analysis of numerous examples in statistical mechanics and field theory, as far as their critical or short-distance behaviour is concerned. These principles are not self-evident, and they do not follow from some underlying theory. In this book, after introducing the basic CFT prmciDles in-Cha^erL-their meaning is explored without vii

viii actually discussing their origin. This is made on purpose since CFT itself is still in a process of evolution. It may not be a good idea to extensively discuss CFT fundamentals when there is a good chance for their being changed or even replaced by more funda­ mental principles. Nevertheless, modern CFT has its own sound basis. Concentrating on this basis, which I believe will survive in time, gives me a chance to represent CFT in a self-contained way. Compared to a QFT description, a CFT description of a given physical system is less informative. Nevertheless, it contains relevant universal information about high-energy scattering or short-distance (critical) behaviour. The main task of CFT is to calculate QFT correlation functions in the limit where conformal symmetry is thought of as being present. An infinite number of physical degrees of freedom presumably contributes to the correlation functions at very high energies which are hardly accessible in experiments, or near a critical point of a statistical system under consideration. Therefore, one should try to understand the conformal physics without solving a general (very complicated) interacting problem by the methods of QFT. The latter are usually of a perturbative nature and, hence, they are very restricted in applications. Instead, CFT is capable of going beyond perturbation theory, and it is just what this book is all about. The idea of conformal symmetry is not the only one which helps. A related idea is the assumption about locality of CFT operators which should form a complete set of local operators satisfying some kind of algebra to be described by the operator product expansions. A good analogue is provided by non-relativistic quantum physics, where a scattering problem can be reduced to solving a non-linear differential equation. In practice, one usually needs only asymptotics of a solution, not the exact one. It may well be quite difficult to find an exact solution. Knowing asymptotics clearly gives only partial information about an exact solution. However, calculating asymptotics is usually a simpler problem, and it is sometimes possible to reformulate the problem in terms of a closed set of equations known as the 'conformal bootstrap'. It may be very difficult to get such a picture by using the tools of QFT, if it is even possible at all. Specific CFT models, as well as solutions to CFT bootstrap equations, may possess additional symmetries. It is one of the tasks of CFT to describe all possible extensions of conformal symmetry. This leads to remarkable connections with various branches of modern mathematics, such as affine Lie and Kac-Moody algebras and 'quantum groups'. It is a physical problem how to identify the relevant operators for a given conformally invariant physical system.


This book intends to provide an introduction to CFT from the first principles, and to reach the current frontiers of research in this area. Throughout the book I tried to keep a balance between introductory and advanced topics, sometimes omitting general proofs and substituting them by examples. This seems to be quite reasonable for describing a theory which has yet to establish itself in a strict mathematical way. Special attention was paid to making a clear distinction between the well-established mathematical results and their applications in CFT. The monograph is organized as follows: Chapter I is devoted to the fundamentals of conformal symmetry in field theory, where a few basic examples are also considered in detail. In Chapter II, I review representations of conformal algebra, conformal boot­ strap equations, fusion rules and the most important particular constructions. This Chapter plays a key role for CFT applications considered in the following Chapters. The partition functions and the Virasoro characters are introduced in Chapter III. The approach I adopted is based on free fields. fields. Affine Lie and Kac-Moody algebras, and their role in CFT are discussed in Chapter IV. Their classification and free field rep­ resentations are reviewed. An underlying topological and group-theoretical structure of the Wess-Zumino-Novikov-Witten theories possessing such symmetries is explained. (Extended) supersymmetry in CFT is discussed in Chapter V. A generalization of the group-theoretical results of Chapter IV to cosets is the topic of Chapter VI. Chapters I-VI essentially comprise the introductory part of the book. Chapters VII-XI contain more advanced material which is close to the frontiers of current research. In particular, Chapter VII is devoted to the W algebras generalizing the conformal Virasoro algebra. In Chapter VIII, I briefly discuss strings and superstrings from a viewpoint of CFT. Two-dimensional quantum gravity and topological field theories in relation to CFT are addressed in Chapter IX. Matrix models in relation to two-dimensional quantum grav­ ity and CFT are reviewed in Chapter X. Chapter XI concludes the odyssey through CFT by addressing some mathematical relations between two-dimensional integrable models and CFT. My attitude towards conformal models as the particular integrable models becomes finally justified in this Chapter. Finally, Chapter XII plays an auxiliary role. There I collected comments about the Literature, mathematical details about Lie and affine (Kac-Moody) algebras, Riemann surfaces, BRST and BFV quantization, and 'quantum groups'. The mathematical comments are written in a pedagogical way, so that they may serve as introductory notes as well. Almost every section contains a few exercises. Most of them are not difficult at all, but they may help for an individual's assessment. Some marginal information used in

X x

the text is put into the exercises labelled by a star (*). For the reader's convenience, the list of abbreviations used in the text is supplied at the end of the book, as well as an index. Key words in the text are emphasized by different type styles. Including all the relevant material for a wide range of CFT topics covered in this book is obviously impossible, and it was not my desire at all. Though being rather extensive, the bibliography collected at the end of the book and my comments about the Literature in section XII. 1 are by no means complete. They only reflect my personal preferences in selecting the material. I apologize to the authors whose papers somehow escaped my attention or were cited inappropriately throughout the book. This book grew out of my lectures given at the Physics Department of Maryland University in 1991-1992. A major part of the selected material is based on existing reviews and original papers. All the editorial work on the manuscript was done at the Institute of Theoretical Physics in Hannover. It is my pleasure to thank colleagues, friends and students, as well as participants of Spring and Summer Schools at the International Center for Theoretical Physics in Trieste over the years, for numerous discussions on CFT and strings. In particular I would like to thank Nathan Berkovits, Jim Gates Jr., Marc Grisaru, Brian Dolan, Marty Halpern, Sergei Kuzenko, Dima Lebedev, Olaf Lechtenfeld, Alexei Morozov, Hermann Nicolai, Hitoshi Nishino, Yaroslav Pugai, Jens Schnittger, John Schwarz, Warren Siegel, Harald Skarke, Arkady Tseytlin, Igor Tyutin and Cumrum Vafa, who helped me to understand some particular issues. I am especially grateful to Olaf Lechtenfeld for his careful reading of the manuscript and critical remarks. The T£X typesetting system combined with the original macros and figures was used by the author in preparing the MpjX file for printing the camera-ready manuscript. The assistance of Jens Johannesson and Thorsten Schwander in choosing the computer software is appreciated. At last but not least, this book would never appear without my wife Tatiana, supporting me and taking care of our children for all these years. Sergei V. Ketov Hannover, Germany October, 1994

Contents I


Conformal Symmetry and Fields



Conformal invariance



Symmetries and currents



Operator product expansion


Central charge and Virasoro algebra


Free bosons and fermions on a plane



Conformal ghost systems


15 . .

Representations of the Virasoro Algebra



II. 1 Fields and states


11.2 Correlation functions


11.3 Conformal bootstrap


11.4 Null states 11.5 Fusion rules 11.6 Coulomb gas picture

. . .



. . . .

55 57

II. 7 Ising and other statistical models . .

. .


II. 8 Bootstrap in the minimal models


II.9 Felder's (BRST) approach to minimal models





Partition Functions and Bosonization 111.1 Free fermions on a torus





111.2 Free bosons on a torus


111.3 Chiral bosonization . .


83 .





111.4 Chiral bosonization on Riemann surfaces . . .

96 .

111.5 BRST approach to minimal models on a torus . IV



. .

. • 104 ...


A K M Algebras and W Z N W Theories


IV. 1 AKM algebras and their representations

. .

IV.2 Sugawara-Sommerfeld construction . . .


IV.3 WZNW theories

. .

. .

IV.4 Free field AKM representations

. .


. .








IV.5 A-D-E classification .


V Superconformal and Super-AKM Symmetries


V.l Superconformal algebras and representations


V.2 Super-AKM algebra and its representations


V.3 Supersymmetric WZNW theories



V.4 Chiral rings and Landau-Ginzburg models VI


. .

Coset Models


VI. 1 Goddard-Kent-Olive construction . . . . VI.2 Kazama-Suzuki construction .

. . .

VI.3 KS construction and LG models VI.4 Gauged WZNW theories .



. .

. .

VI.6 Generalized affine-Virasoro construction



175 179


VI.5 Felder's (BRST) approach to coset models

. .

183 .

. ...

. 189 195



W Algebras


VII. 1 W3 algebra and its generalizations


VII.2 Free field approach


VII.3 Quantum Drinfeld-Sokolov reduction


VII.4 W coset construction .


. .

VIII Conformal Field Theory and Strings

227 231

VIII.l Bosonic strings in D < 26


VIII.2 Supersymmetry and picture-changing


VIII.3 Extended fermionic strings


VIII.4 W gravity and W strings . . .


. .


IX 2d Gravity, and Topological Theories


IX.l Quantum 2d gravity


IX.2 Liouville theory


IX.3 Topological field theories and strings


. . . 307

X CFT and Matrix Models


X.l Why matrix models ? . .


X.2 One-matrix model solution X.3 Multi-matrix models





X.4 Matrix models and KdV


XI CFT and Integrable Models


XI. 1 KdV-type hierarchies and


. . . .

. . . .

XI.2 W-flows


XI.3 WZNW models and Toda field theories XI.4 4d self-duality and 2d integrable models XI.5 Self-duality and supersymmetry .



.... . .

363 371

. . 374

xiv XII



XII.1 About the Literature


XII.2 Basic facts about Lie and AKM algebras


XII.3 Basic facts about t9-functions

. .

XII.4 Basic facts about Riemann surfaces . . XII.5 Basic Facts about BRST and BFV XII.6 CFT and quantum groups



. .



. ■



419 424



Text Abbreviations




Chapter I Conformal Symmetry and Fields In this Chapter we introduce basic concepts of conformal field theory (CFT), study basic examples of two-dimensional (2d) conformal fields and calculate their correlation functions on a plane; The operator product expansion and the conformal algebra are shown to play key roles in CFT.


Conformal invariance

The basic ingredients of any field theory are a set of fields $ A (z) depending on spacetime coordinates zM, and the action S[$] to be a functional of the fields. In quantum field theory (QFT), one usually starts with the field theory action and then one calculates the correlation functions via the path integral approach. Our goal will be to calculate the correlation functions in CFT directly, by exploiting its symmetries. In this process, the action itself becomes redundant. In fact, more than one action may correspond to the same physical theory. A field theory is normally classified by some set A of fundamental fields, which are generically infinite in number. The fundamental fields are supposed to form a basis in the field space. Though space-time is usually assumed to be equipped with the Minkowski signature, we choose to work in Euclidean space, so that our spacetime vector indices will be represented by the middle Greek letters, /A, V, . . . = 1,2,..., d. We find it convenient to allow the number d of Euclidean dimensions be arbitrary at the beginning, since it helps to understand the special role of two dimensions in CFT. 1




Our consider arbitrary arbitrary field field theories theories (there (there is is not not much much Our goal goal is is not not just just to to consider about them) but those which are reparametrization-invariant. This means about them) but those which are reparametrization-invariant. This means that that the the fields fields of of the the theory theory there there is is aa d-dimensional d-dimensional metric metric ('gravity') ('gravity') g^{x), g^{x), and and the the is is invariant invariant under under the the transformations transformations

to say say to among among theory theory

5x» fa" = £"(x) e"(a;) , WA A (( II )) = = iL£e$&(x) W * 4 W ,, x £ Sg^u(x) d\g^„ Sg^(x) = Qp£u dp£v ++d»d„e„ " ~-£xed\g» v , ,


where eti{x) are the infinitesimal parameters, and LE is the Lie derivative. The fields $ A are supposed to transform according to their tensor nature. The metric g^x) defines an invariant line element ds2 = g^dx^dx", and it transforms under a finite transformation i ' -> i*1 as a rank-2 symmetric tensor dxx dxf dxp 9^{x) g^(x) -> -4 g^ix) 9 {i) = -Q^-Q=p9\e(x) ~ »" = d ^ W 9 ^


The invariance under reparametrizations is not in fact even a restriction, since any field theory is usually assumed to be either Lorentz-invariant (in a flat space-time) or general coordinate invariant (in a curved space-time), and this implies some metricdependence in the theory. 1 We assume almost everywhere in the book that our Eu­ clidean spacetime is flat and isomorphic to R d , so that there exists a globally defined coordinate system in which g^ = 5^. 2 After being quantized, the fields M(ZM)) 1


= = 77 vv ''-- 1/ /nn[ [^^AA] ]**AA1 1( (xx1 1) $) $AA2 2f fee) -) - - *- *AAMM( x( xAAf f) e) e- - lS^l l^ l ,,

'This is not true for the topological field theories to be considered in sect. IX.3. What happens in the presence of quantum 2d gravity is discussed in Chs. IX and X.


(1.3) (1-3)



where the functional measure is yet to be specified, and N is the normalization factor, N = I U[D$&] e-W^. The VEV (1.3) is called a correlator or a correlation func­ tion. The quantum field theory is solved once all its correlators are calculated. One of the standard ways of doing that is to use the perturbation theory (the Wick's theorem) usually described in terms of the conventional Feynman graphs. Quite generally, there is no way to exactly calculate the correlators, unless there are powerful symmetries in the theory. The general idea is to impose such symmetries on QFT, which would severely constrain the correlation functions in such a way that they could be calculated, while being non-trivial. The CFT's are a subclass of QFT's, which are invariant under the conformal transformations. It is of interest to test the power of the conformal symmetry in quantum field theory, and to explore the consequences of this invariance. This will be the main story in our discussion. The physical examples where the conformal symmetry is particularly relevant are represented by strings in high-energy physics, and critical phenomena at the second order phase transitions in statistical physics. By definition, the conformal transformations are the restricted general coordinate transformations (diffeomorphisms), x —> x, for which the metric is invariant up to a scale factor, (l) (1.4) gAx) -> ~> 9^(x) = 0(*)fl^ ^l(x)glu, , 0(x) f2(x) = ewu . g,»(x) foffl = Clearly, these transformations form a group, which is known as the conformal group. Since our space is flat, the reference metric can always be chosen to be flat also, g^v = 8^. = SpUOur first task is to identify the conformal transformations from eq. (1.4). They Bj\/A2B2 A■BB == clearly preserve the angle A ■ B/VA between any two vectors A1* and B^, Ag^A^B" That's why such transformations are called conformal. They also contain the Poincare transformations (translations and Euclidean rotations of the flat space Rd) as a subgroup, since the latter obviously satisfy eq. (1.4) with g^ = 5 x^ + e>1(x) w.r.t. the flat metric g^ = 5^. One easily finds from eqs. (1.2) and (1.4) that 22 (1.5) Sp, ++ dd^Sv w x llev + dveli = Sl5lu, = [1 + u(x)]6ia/ , Sfw + 3 ^ = 0,5^ = [1 + w(as)]V . w(x) ( ) == - J9 ^%^ .■ Hence, one gets the equation 2 e„ + top ==| -j(9 ( S •«) 2. SubstiSubsti­ From tuting the general second-order polynomial in x into eq. (1.6) and fixing its coefficients by that equation yield the result that the conformal algebra consists of the ordinary 1 1 M translations (e* ') and rotations (e i ^ x " , ^ , , = —u^), scale transformations {e^ = = aa?) (e*1 = aui^x"^^ i ,t (e' = \x^) Ax") and the special conformal transformations (e* (e*4 = = fx2 — 2x Ix^b spe­ b • x). AAspecial conformal transformation can be recognized as a composition of an inversion and a translation: x^/x22 = x,1,1jx22+b,i,i. The whole algebra is locally isomorphic to so(d+l, 1). 33 The conformal group is comprised of finite conformal transformations. They include the Poincare group -> x" i " = x" + a" , x" -4 x" -4 St* x" ==A"„x" x" A^x" , , A^ A", e6 50(d) 5 0 ( d ),,

(1.8) (1.8)

and, in addition, the scale transformations (dilatations) and the special conformal trans­ formations, resp., i " -4 -»• 5? i " = Ax" of As" , w = - 2 1 n A , Tu

, Z.V. _

2 WT rf J-4--rWT u x



2 2 (1.9) 2 2 "^ l 1++ 2b (1.9) 2ft •xx+ +b6x2x' 2 - = 21n[ll + 26.x + 6 V ] J It is important to realize that the d-dimensional conformal group is finite-dimensional for d > 2, and all its transformations are globally defined. It has been known for a long time that the conformal group is a symmetry of massless fields with dimensionless coupling constants, and that it is the maximal kinematical extension of relativistic invariance [5].


i , i =

(1.7) . When The case of two dimensions is special, as one can already see from eq. (1.7). 9iw = S,u, and d = 2, eq. (1.6) takes the form of the Cauchy-Riemann equation diE diEXX = d22££22 ,

dx£ £i . d\£22 = -d22£i


In terms of the complex coordinates and fields + ix2 , z = x11 + {z, z) = el(z, z) + ie2(z, {z, S) z) , ez(z,

z = xl1 — ix2 , (z, z) = es1l(z, z) - ie2{z, iezz(z, (z, z) , z



(1.11) z

= £ {z) and ei = i*(z). i (z). Thereresp., eq. (1.10) implies a holomorphic dependence: ze = There­ fore, the 2d conformal transformations can be identified with the analytic coordinate transformations: ~t f(z) z -4 f(z) , f'(z) # 0 . z -> f{z) (1.12) 3

Given a flat d-dimensional space with an indefinite metric of signature (p, q),p + q = d>2 corresponding conformal algebra is isomorphic to so(p + 1, q + 1) [5],





This can be seen in yet another way, when rewriting the line element in the complex coordinates on a plane as ds2 = dzdz. Under the holomorphic transformations (1.12), we we get gei



dzdz -> —> -J- dzdz , (1.13) dz dz which just means that we have a conformal transformation, in accordance to the definiwhich just means that we have a conformal transformation, in accordance to the definition (1.4), with u(x) = 21n|d//dz|. tion (1.4), with u(x) = 21n|d//dz|. It It is is possible possible to to define define the the infinitesimal infinitesimal analytic analytic functions, functions, parametrizing parametrizing the the conforconformal mal transformations, transformations, by by restricting restricting the the plane plane to to aa finite finite region region around around the the origin origin and and assuming that all the singularities of the analytic functions are outside the region chosen. assuming that all the singularities of the analytic functions are outside the region chosen. Since Since eq. eq. (1.10) (1.10) allows allows an an arbitrary arbitrary dependence dependence on on z, z, it it is is quite quite reasonable reasonable to to assume assume this this dependence dependence to to be be meromorphic meromorphic in in general. general. + e(z), e(z), The suitable basis for the infinitesimal conformal transformations z —¥ z = z + + e{z) is given by and z —> 1 = z + n+1 1 +1 1 n+l -+ zz* -- aannzzn+1 n 6€ ZZ ,, z* -+ ,, zH-> iz -- ao^a„z" nz " ,, n

(1.14) (1.14)

which are generated by the operators lkn = = -zZn+l —In , ln == -z2n+1 — , dz dz

~ 'ah'

~ fz'

neZ n6 L .


(L15) (1.15)

These operators satisfy an algebra which consists of the two commuting pieces, (n-— m)ln+m , [l [In,n, lm] = (n - m)/~ im] = (n [in, lm] m)/ nn+m = 0, + m , [/„, lm] =


each one being known as the 2d local conformal algebra. The independence of the two algebras {/„} and {ln} justifies the use of z and z as the independent coordinates. This literally means a complexification of the initial space: C —» C 2 , which gives us a freedom to choose various reality conditions by making 'sections' in the complexified space. In particular, the section defined by z = z" recovers a Euclidean plane. A Minkowski plane can be recovered by the section z* = —z, which implies (z,z)z) = i(f + (z,

ff,f-ff) ff,f-ff), ,

=-df da2 2. ds ds2 2 = = -df22 + + da dd


However, these are not the coordinates we want to associate with the 2d Minkowski = ec, z = = e c , with Q = space-time. First, we make the conformal transformation z = T + ic , £ = T — iff, from the z-plane to a cylinder, —oo < r < +oo, 0 < a < 2ir. By




definition, the Minkowski space-time formulation of a field theory is obtained from its Euclidean formulation on the cylinder by the Wick rotation, C = r+ia -+ i(r+a) = iC + , £ = T — ia —► i(r — a) = iC,~, where the Minkowski light-cone coordinates £* = r ± a have been introduced. In terms of the Minkowski light-cone coordinates, the line element takes the form ds2^ = d(+d(,~, and the conformal transformations take the form of the reparametrizations of C+ and (~ : C'+ = /(C + ) . C'~ = s(C~)i which leave the light-cone invariant. The line element ds2^ is clearly preserved by these transformations up to the scale factor ft = /'(C + )s'(C~)The relation between Euclidean and Minkowski space-time formulations of CFT is important, to understand the quantization in what folllows. We will frequently discuss the holomorphic dependence only, and ignore the similar anti-holomorphic dependence. The number of generators in the 2d local conformal algebra (1.16) is infinite. Obvi­ ously, the infinite number of symmetries should imply severe restrictions on the conformally invariant field theories in two dimensions, and that's what our story is all about. It is also worthy to mention that the conformal transformations (1.16) we consider are neither globally well-defined nor invertible, even on the Riemann sphere S2 = C U co. This is because the vector fields V(z) = - 5 > n / n = £ a n *z " + 1 - ^ , V(2) = z n




generating holomorphic transformations, are globally defined only if an = 0 for n < —1 and n > 1. This is necessary for the absence of singularities of V(z) at z —> 0 and z —► co (use the conformal transformation z = — X/w in the latter case!). Therefore, a global group of the well-defined and invertible conformal transformations on the Riemann sphere is generated by {l-\,k,k} U {1-iJa,k} only. Eq.(1.15) tells us that /_! and L a can be identified with the generators of translations, (Z0 + IQ) and i(/ 0 — IQ) — with the generators of dilatations and rotations, resp., and k and k — with the generators of special conformal transformations. The corresponding finite transformations form a group known as the complex Mobius group : az + b # = — , (1.19) 1.19 zz-¥z -+ z = v(1.19) cz-Yd , ' cz + d where a, 6, c, d are complex parameters and ad - be = 1. The group of transformations (1.19) parametrized by six real parameters is isomorphic to SL(2,C)/Z2 = 50(3,1). = The need for the quotient Z 2 is caused by the fact that the transformation (1.19) is not sensitive to a simultaneous change of sign for all parameters a, b, c, d. This group is sometimes referred to as the group of projective conformal transformations.




The lesson we learn from here is that the local conformal algebra in two dimensions cannot be fully integrated to a globally defined group, but it is the full local conformal symmetry (1.16) that will be the symmetry we want to investigate. Representations of the global conformal algebra (after quantization) assign quantum numbers to physical states. It is quite natural to assume the existence of the vacuum state |0) among the physical states, which is invariant under the Mobius transformations and has vanishing quantum numbers. The eigenvalues h and h of the operators Z0 and l0, resp., are known as conformal weights of a state. Given the conformal weights of a state, its scaling dimension A and spin s are determined by A = h + h and s = h — h, in accordance to the similar assignment for the generators of dilatations and rotations. Instead of the Riemann sphere, which is a closed Riemann surface of genus zero, the simplest open Riemann surface to be represented by an upper half plane (with the infinity attached) could equally be considered. In this case, the parameters a, 6, c, d in eq. (1.19) have to be real, thus leading to the real Mobius group SL(2, R)/Z2- There is an obvious similarity between the 'open' case and a holomorphic part of the 'closed' case.

Exercises # 1-1 > Consider eqs. (1.6) and (1.7) and show that the third derivatives of e(x) must vanish when d > 2. # 1-2 > Calculate the number of generators of the conformal algebra in d > 2 dimen­ sions and compare it with the dimension of the Lie algebra so(d + 1,1). # 1-3 > (*) Rewrite the two-dimensional conformal transformations (1.8) and (1.9) in terms of the SL{2, C) matrices. What is the well-known group theory decomposition which is illustrated by the result?


Symmetries and currents

Let us go back to a general field theory defined in an arbitrary number d > 2 of Euclidean dimensions. The reparametrization symmetry (1.1), like any other symmetry of the field theory action 5, implies the Ward identity for its correlation functions defined by




eq. (1.3). This identity follows from the obvious equations: m

£ ($A,(zi) • ■■t$Aj(Xj) ■ ••9&„(XM)) 3=1 m m


5 | a5!| t i l *A, (*I) • • -»^(*i) •" ■A *A = N~l J/ nIIP>*A] P>«A] £ £;)---*A ,(zM) M (%) e - e*/ i=\ J'=I

= \ J d2x^g8gml/(x)

( T " ' ( * ) * A I ( * I ) ■ • ■ * A M («*)) ,


where the stress-energy tensor T'"/ of the field theory has been introduced, T»u,x)





s/gSg^x) y/gSg^ v(x)

In the flat This tensor encodes the reaction of field theory to the metric deformations. deformations. In space with the reference metric g^v = 5^, eqs. (1.1) and (1.20) imply m


({«A,(»I) $A,(SI)

■■5$cji(xj) (xj) ■ -- *^A1„ % • ■■s$* ( %)) ))



=yjdd2xd^(x) = xd^(x) vwi^fa) VWi^xi)

■ ^ M(IM)) ) , ■ -■t •$ 4M



which is just the reparametrizational Ward identity between the correlation functions we are looking for, since the stress-energy tensor itself is a composite operator in terms of the fundamental (basis) fields. The Ward identity (1.22) is quite general, and it is valid for any reparametrization invariant field theory. It simply means that the stress-energy tensor components are the generators of the general coordinate transformations. In particular, in a flat space, the T^ themselves generate translations, whereas the x^T^ — XxTw are the generators of Euclidean rotations. These currents are both conserved, 3 T F = 0. The latter follows from eq. (122) after integrating by parts and using the arbitrariness of the parameter, and it is clearly valid for all points on the plane except the distinguished ones (called punctures) {xt}, where the field operators are inserted. We actually want to study the conformal field theories having a larger symmetry then the Poincare group. Let us consider first the consequences of the global scaling invariance x» -> Ax*1, which is a part of the global conformal group. The associated current is given by J„ Jfi = x"T * ■*vvk Vfl . •





The conservation of the current (1.23) is equivalent to a tracelessness condition for the stress-energy tensor, T 7 ^ = 0. This restriction is known as the scale invariance condition. It implies, in particular, the vanishing of any correlator between $A'S and T*. It is easy to see that any current of the form fx(x)Txv will also be conserved provided dT #7" + + dT d"f» - ip{x)rT 4 4>a (ZM,ZMZ))M)) J2 (* ( * A"" (( 2 li^l) l i ^ l ) •• ••• •^AjiZj, (ZM, MM 2 = J z zedzz{z,z){T e*(z, z) (T Z 2 (z)$ z{\M,z ■M■ jd2dzd)) ■ $ A M ( Z M , ZM)) , , A l fa,M{z Z2{z)^^{z uz,)---^^


where only the variations w.r.t. the holomorphic coordinates were considered. A quite similar formula holds for the variations of the anti-holomorphic coordinates. We thus effectively reduced the 2d dependence of CFT to a 1-dimensional dependence: (z, z) —► z. This observation is very crucial for the whole integrability of CFT, as we shall see all the time throughout the book. The generators of infinitesimal conformal transformations can be defined in terms of T(z) as

n+1T W ^ ^ with the contour encircling the origin. The formal operatorial equation (1.32) makes (1.32)

actual sense when acting on fields whose arguments are inside the integration contour. The contour shape is irrelevant because of Cauchy's theorem. The same theorem yields the statement n 2 L L z(1.33) T(z) = ■£ T{z)=Y, ^-n n-\~ , neZ neZ

and similarly for the Ln and f. We now have all the power of complex calculus at our disposal. To incorporate the standard machinery of canonical quantization into CFT, it is convenient to use the radial quantization techniques [22, 23, 24] on a plane. It uses the following parameterization of C 2 : e c, z = e^

£C,==r T++iaia, ,

(1.34) (1.34)

in terms of the Euclidean 'world-sheet' time and space coordinates, r e R and 0 < a < 2ir, resp.


5 Eq. (1.34) is just the definition of the r and a used here. In string theory (Ch. VIII), they can be interpreted as the coordinates of the Euclidean closed string world-sheet. This gives the useful vizualization of quantization procedure.




Fig. 1. The conformal map of a cylinder to a plane. We interpret eq. (1.34) as the conformal map of a cylinder to a plane (Fig. 1). Infinite past and future, r = =FOO, on the cylinder are mapped into the points z = 0, oo, resp., on the plane. Equal-time slices are the circles of fixed radius on the plane, whereas equalspace slices are the lines radiating from the origin. The time translations r —> r + X are the dilatations on C: z —> —> eexxzz = >►a a+ + 66 = zz + + Xz Xz + + ..., ..., whereas whereas the the space space translations translations aa — — are the rotations on C: z —> e'9z. Therefore, the Hamiltonian of the system can be identified with the dilatation generator on the plane, while the Hilbert space of states comprises surfaces of constant radius. The stress-tensor components T(z) and T(z) in eq. (1.30) are identified with the generators of local conformal transformations on the z-plane. In the radial quantization, an 'equal-time' surface becomes a contour on the z-plane surrounding the origin. To convert these heuristic arguments into precise statements, consider the time evo­ lution of any operator $ in the Heisenberg picture with a Hamiltonian H. The dynamics is described by the operator equation of motion

£^ -=[ *[ #•. *]]..

(1.35) (1-35)

where the square brackets mean an equal-time commutator. Equivalently, eq. (1.35) can be rewritten to the infinitesimal form f i zT(z) = L00 , z + e(z), one should take


L37 (1.37)

< >




as the conserved charge in the equation of motion -M>(a)

(1.40) d.40)

where the conformal weight (h, h) is real-valued. The rest of CFT fields are called secondary fields. The infinitesimal form of eq. (1.40) is ${z, z) = [{hds 5eJl [(hde ++ ed) ed)++ (Me (hde++id)] id)]*(e, *(e,i)i) , , e^{z,


where the shortened notation has been introduced, d = dzdz,

BB = d-2.


Eq. (1.40) is useful to relate mode expansions of a primary field on the cylinder and on the plane. Taking, for example, a holomorphic field with h = 0, its Fourier expansion on the cylinder, (1.43) £*■ " *C .■ CC ==rr ++ iff, ia, (1.43) *A( 0 = E ^ e«"" can be transformed into an expansion on the plane, n- nh- / i « f c W == Yi £ Kz~ *kW nz~ ~ .


n€Z nCZ

Eq. (1.33) is now recognized as the particular case of eq. (1.44). This means that the stress tensor has conformal dimension h = 2.




Let us now consider the constraints imposed by the conformal invariance on the correlation functions of the primary fields in a quantized 2d field theory. This means that its correlation functions do not change under the transformations (1.41). The two(2) {z (z22, ,zz22)) point function G(2) (zj, 2j) = ($i(zi,z ($i(zj, Zi)$ )) provides the simplest example, where {,Zj) 1 )* 2(z an invariance under conformal transformations leads to a partial differential equation of the form l(e(zi)d hide(zi)) + (e(z2)dZ2 + M [{e{z ( z 22)))) [(e(zi)d h2ede(z x)dzi zizi + hdeizi)) -■00. . hi§e(zi)) + {e(z )dZi = hide(zi)) (e(z22)d, de{z22))] G™{z G^'fe.zV) +(e(zi)d-zizl + Si&(%)) + h2§e(z 2 + Z2 %, z\)==

(1.45) (1.45)


Zi] 22 = = Substituting e(z) = e(z) = 1 implies that G depends on zi2 = zzxx -- z22 and z (2) 22 +h2+ h 2+ h2+ S22 z f ^ ) zi-z \ z ; e(z) = G< G ' G Gi /(z^ '' z^ zzi -z%\ substituting e{z) = z and e(z) = z leads to G< > = C /(z^ z^ ' ) ; finally 12 x 2 2 l2 substituting e(z) e(z) = = h2 = h , hi = hh22 = = h (otherwise = z22 and e(z) = z22 gives hi = G(2) = 0). Putting all together determines the form of the two-point function as (1.46) (1-46)

G ( 2(zi,z,) C f e ) -= ;ZZz §ZZz . h! ,, 1 222 1 2

where the G12 is a constant to be fixed by normalizations of the fields. $ i $ 2 $ 3 > is determined along the similar lines, The 3-point function G' 3 ' = < $i$2$3 G G

(Z„ Z,) = - G l11222333 hl+ h2+h Zi) = ((Zi, Z j , Zi) ,++2h-h h2 j_3- hh a, Ju+hz-hi _h +fc33-hi -fci ^ 3 h +3' +hi-h > l - ' > 22 h1], h Z12 Z23 Z13 Zj2 Zj3



--hi+h /IJ+H -h3 2 - S 3 J=h,2+h.3-hi-h.3+hi—h i 2 + h 3 - f c l z =h 3+hi-h22 ^KS+K-L-KI z13 13

fl 471 (1.47) /


2 z-h-i+h2-h-i-h2+K3-hi z z l 2 z23 12 23

V • V •


Gi23 is a constant. Hence, we can conclude that the coordinate where zy zy = = Zj z* — — z.,, Zj, and and G123 GIM where is a constant. Hence, we can conclude that the coordinate conformal invariance. The dependence of the 2and 3-point functions are are determined determined by by conformal conformal dependence of the 2- and 3-point functions invariance. The , z can be mapped geometrical reason why it happens is because any three points geometrical reason why it happens is because any three points zi,z zi,z22,z ,z33 3 can can be mapped by a complex Mobius transformation into any three reference points, say 00,1,0, where by a complex Mobius transformation into any three reference points, say 00,1,0, where 2 lira lim zz22hAl zl zf'G< ^ G((333))>=G = G11221332 3 ..■ ^G =G


The coefficients like C123 carry a nontrivial dynamical information about the operator algebra of a given CFT. Knowledge of these coefficients is equivalent to solving CFT. The equations satisfied by these coefficients are discussed in Ch. II. Things become more complicated for the n-point correlation functions with n > 4. The global conformal invariance restricts the 4-point function of the primary fields to the form [17]

)+h/3 )+h/3 G» (Z„ ^=/(*, = / ( * , x) x ) nn /z/^ n * /n z-f^ *'" -,






(1.49) (1-49)



15 15

where h = £f_j h{ , h = £ Derive eq. (1.24) and find the explicit form of functions fx{x) and ip(x) for the case of the special conformal transformations. # 1-5 t> Calculate the two-point correlation function of quasi-primary fields in d dimensions from conformal invariance. In d > 2 dimensions, the higher-point correlation functions cannot be fixed by the conformal symmetry alone.


Operator product expansion

To be well-defined, eq. (1.38) requires a resolution of the operator ordering ambiguity. The problem arises in the radial quantization because it uses the 'equal-time' commutators at \z\ = |iw|. In general, an operatorial product A(z)B(w) on a Euclidean plane is defined only for \z\ > \w\, since this is equivalent to the time-ordering in QFT. Therefore, we can proceed in the usual way known there to resolve the operator ordering problem, namely, by using Schwinger's time-splitting technique [1]. The latter implies 6 efinition: 6 the following definition: lim { UIi ¥-£{z)7l(T(z)${w. [T„ [TE, * *(«,, K w)] ~ . e(z)K (T(z)*(w, *)) iB)] *M lim (T(z)*(v,, *)) l~*M |y|z|>| \z\->\w\ yj\z\>\w\ |z|->M m| 2m 2m dz

^s(z)7l(T(z)$(w,w))} e(z)7l(T(z)$(w,w))\ ^s(z)K(T(z)$(w,w))\ £-Mz)K(T(z)*(w,iB))\ -i7|z|| 7|z|| 2m 2m






/|z|M> I ^ ! " ! '

The anti-holomorphic part is assumed to be added everywhere, when it is necessary. We normally drop the 7^symbol and always assume that the Euclidean correlators are radiallyordered in what follows. 7




Fig. 2. For the evaluation of an equal-time commutator on a conformal plane.


r dz = lim f — —e(z)T{z)9(w,w) e(z)T{z)$(w,w)




where the contour encircles the point w in the z-plane. This integral is clearly nonvanishing only if there is a singularity in the operator product \imT(z)$(w,w) lim T(z)$(w, w) ,



whose residue contributes to the r.h.s. of eq. (1.52) by Cauchy's theorem theorem. In general, when two operators A(z) and B(w) approach each other, Wilson's oper­ ator product expansion (OPE) [25, 26] takes the form A(z)B{w) - ™ w)0 A(z)B(w) ~ £ J2CC ^ZA(z ~ ) OA(w) A W .,



where the {OA(W)} are a complete set of local operators, and the CA'S CA'S are are (singular) numerical coefficients. Equations like eq. (1.54) have to be understood as being valid when the product A(z)B(w) is inserted into a Green's function with other elementary operators of the theory, i.e. lim 111 I A(z)B(w) - £ CA(z - w)0A(w) Jj ^(w *„(w =0 Jim $L(W • • •■$■M(w x)l ) ■ M)\u)\ =■ 0

(1.55) (1.55)

In CFT one can use eqs. (1.38) and (1.41) to introduce the OPE between the T(z) and a primary field $(w, w) in the form T(z)$(w, w) rr*(w, dwd$(w, T(z)(w, w)== -. rr*(w,w)w)H-\ w) w$(w, w) (z — u>y z z—— ww (z — u>y 2)

( i)

+ $ ( _ 2 ) {w,w) ( M ; , W) + iu)$(-"3)(w,w) (t(;, w) + ... ,, +&+ (Z(z-w)





where the dots represent an infinite set of other regular terms depending on the new local fields called the descendants or the secondary fields w.r.t. a primary field $ of conformal dimension h. The secondary fields are determined from the above equation (1.56) by the contour integration: &~n)n)(w,w) $*(™) , w (z (z — — wY w)z zz — —VJw keeping keeping only only singular singular terms terms on on the the r.h.s., r.h.s., and and displaying, displaying, as as aa rule, rule, only only the the holomorphic holomorphic dependence. Only these terms are relevant for the correlation functions of dependence. Only these terms are relevant for the correlation functions of the the primary primary fields with the stress tensor. All the operatorial relations we are going fields with the stress tensor. All the operatorial relations we are going to to consider, consider, and and the the OPE's OPE's in in particular, particular, should should be be understood understood as as being being inserted inserted inside inside correlation correlation functions. functions. The The latter latter carry carry all all physical physical information information about about CFT, CFT, and and they they are are therefore therefore the only ones we need to know. the only ones we need to know. The primary fields {$;} in CFT are conveniently normalized by taking their 2-point correlation functions to be of the form l = ^S{zij{z _\)2k («,(*, z* ) $ > , w))U)= _l_w)2hi{2 _\ s )2k w)2hi{2

-S«-"0 * (*-*)*


The OPE for a product of two primary fields takes the form s s K s z w hk hi hj fht 4 hi hhk $i(z,z)$j(w,w) (*) Derive eq. (1.80) from the Jacobi identity for the commutators. # 1-11 > Use the group property (composition law) for the infinitesimal conformal transformation (1.81) of the stress tensor to derive its finite form in eq. (1.82). The Schwartzian derivative satisfies the following composition law S[z;w] = (dwf)2S[z; S[z;f) f] + S[f; S{f;w}. S[z; w] = w] .


The Schwartzian derivative is the unique object of dimension two that vanishes when restricted to the global conformal group SL{2, C).


Free bosons and fermions on a plane

The simplest example of CFT and the best illustration of the formal aspects developed so far are provided by the two-dimensional theory of a single free massless scalar boson



whose action is 55

a$9$) ,

* == ^^ / ( aa $$ 55 $$ ) '

((1.88) (LL8888))

where the integration measure on a complex comple x plane has been defined in eq. (172). Eq. (1.86) determines the propagator of the (z,z) 2) field, ($(z,z)${w,w)) = - l n | 2z - w | 2 . ($(2,2)$(TO,W)))


The general solution to the classical equation of motion splits the scalar field $ into a sum of holomorphic and anti-holomorphic pieces, $(2, z) 2) = (j>{z) + j>(z) 4>{z). .

(1.90) (1.90)

These pieces are related by complex conjugation in the Euclidean case (that's why one uses complexification of the plane), but independent in the Minkowski case where they become left-movers and right-movers, resp. The decomposition (1.90) is unique modulo constants. The propagators of the (anti)-holomorphic scalar fields read (cj>{z)(j>{w)) = - ln(2 ln(z - w) ,

({z)4>{w)) = - l n ( z - u i )


Though the field and T by using the Wick theorem:



z — W


9 The normal ordering used here is just the radial ordering defined by OPE, or, equivalently, the normal lormal ordering via the Wick theorem in terms of modes. '"Because of the normalization convention we use for the 2d actions, the factor of 2IT should be inserted tiserted into the r.h.s. of the definition of the stress tensor in eq, (1.21).





This means that d(j> is a primary field of dimensionft. ft = 1, and it is now quite legitimate to define its mode expansion as 1 11n1 l id{z)

_._ °di>) W ) ,, ^ = _~ 4Z/ / ( (
J((S: )) ++w W- *- > J (( ;; )) .,

27 (1.100) ™>

where the Pauli matrices (a (tr*, ays, az) have been introduced, with the az playing the role x, cr of the two-dimensional '75'. The stress tensor components for the action (1.99) result in T+[z) T+(z) =\ : ^(z)dxP(z) : , _ _ fj(z) j>{z)8j>{z) :: . . Tj(z) = = || :: 1>{z)d${z)


It is straightforward to derive the OPE's ip(z)il>(w) ip(z)rl>(w)

— , z— —w


and Ti(z)1>(w) ~ . 1 / 2 .Mw) + - i - a V - H • (1-103) (1.103) \Z — W)* Z— W \z — wy z —w Hence, the ip (or tp) is a primary field of conformal weight (1/2,0) (or (0,1/2)). On a cylinder, the MW fermions can have either periodic or anti-■periodic anti-periodic boundary = 1/2 for them, a periodic (anti-periodic) fermion on the cylinder conditions. Since h = becomes an anti-periodic (periodic) fermion on the plane: NS:

i}(e2niz) = + i>(z) V-(z)

i)(e2ni2l,iz)z) = = - i>(z) R : rP(e i/>(z). .


A presence of the two types of fermionic boundary conditions (called the NeveuSchwarz-type (NS) and Ramond-type (R), resp.) is ultimately related to the fact that spinors naturally live on the double covering of a punctured plane. 12 The associated mode expansions for the fermions are NS:

# ( *z ) =


12 rpnz-"' V n z - n -' 1 / 2 ,

n€Z+l/2 6Z+1/2


1/2 n 1/2 R: ^-"" ,, R : #(*)= #(*)= £ £ 4> nz- -



so that the Ramond-type ip(z) V>(z) has a square root cut from 0 to 00. The OPE's given above can equally be considered as just another equivalent representation of the quantum 12

The conformally invariant objects of the type ip(z)(dz)n+l/1 with integer n naturally define spinors. The consistent sign choice for the root implies that the space where they live has to be a spin manifold.




commutators between the operatorial modes and vice versa. In particular, for free bosons and fermions, one has

idA(z) =J2nZ-"(z)=£t/-nZ-"-1/2.


Eq. (1.106) can be inverted to produce an = {z) , J 2iri

107) (1.107)


92) and where all the contours are supposed to enclose the origin. The use of eqs. (1. (192) (1.102) reproduces the canonical commutation rules [a*, Oim] [o»i Om] =n5 =nSn+mfi n+mfl ,

(1.108) (1.108)

{ipn, {ipn, Tpm) Tpm) =(w))KR s= (0| a(oo)i>{z)f(w)a{o) a(oo)i>(z)iP(w)a(o) |0)


to represent the two-point function in the anti-periodic case. The evaluation of eq. (1.110) goes in a straightforward way, when using eq. (1.108) and taking into account the presence of a zero mode i/>0 (ip% = 5) : J CO


- (i>(z)i>(w))K = = ( £ a,*—1" £ \n=0 \n=0



i>mw-™-vA / p_





w + = t *-- »- +\4==~ -1*■*■**-*+\ A-- i- ((—+i) + i) -=^M^!>. v ? ) . (Lin-) (1.llla)






The propagator (1.111) has the same short distance behaviour as the one in the periodic The propagator (1.111) has the same short distance behaviour as the one in the periodic case, and it also changes sign when either z oi w makes a loop around 0 or oo. The case, and also changes when either with z oi w makes a loop[17]. around 0 or oo. The r.h.s. of (1.111a) is thesign unique function such properties r.h.s. of eq. (1.111a) is the unique function with such properties [17]. The conformal dimension ha of the twist field a(w) can be extracted from its OPE The conformal dimension ha of the twist field a(w) can be extracted from its OPE with the stress tensor, and it turns out to be (exercise # 1-15) ha = 1/16 [33, 34, 35, 36]. with the stress tensor, and it turns out to be (exercise # 1-15) ha = 1/16 [33, 34, 35, 36]. The bosonic twist field could also be introduced by defining the two-point function as [17] (d4>{z)d(j>(w))R = (0| a(oo)d4>{z)d(j>(w)a(0) |0) |0>

= -_ | l((zf(—+^w) f| ))-




The twist field a(w) has dimension ha = 1/16 and satisfies the equation 1/21/2 d(j>(z)a(w) T(W) d(z)o(w)~ ~(z(z- -W)~ W)T(W) , ,

(1-112) (1-112)

where the new twist field r of dimension hT = ha + 1/2 has been introduced [17]. A nice intuitive picture can be provided for treating the twist fields, where a cut along which fermions change sign is equivalent to an 50(2) gauge field concentrated along the cut (the Dirac string singularity). The gauge field strength can then be adjusted to give a phase change n for a parallel transport around the endpoints of the cut. Given this picture, the twist field looks like a point magnetic vortex, whereas changing the position of the cut corresponds to a gauge transformation of the 50(2) gaugefield[17]. Exercises # 1-12 > Perform a detailed calculation leading to eq. (1.96). # 1-13 t> Check eq. (1.98). # 1-14 > Derive eq. (1.108) from eqs. (1.92) and (1.102). # 1-15 > (*) Determine the conformal dimension of the twist field a by evaluating the expectation value of the stress tensor (i.e. the leading term in the OPE of T with a): l tTlrW I f) 1 (T(z)) (dw!i'i(-}rfll,, ,,


-ib(z) = ££& &„„*2 - " - *A -;&(*)



T+ r To quantize this theory, one uses the parametrization z = eT+t "",, and rr as the evoluevolu­ tion parameter. This yields the canonical equal-time anticommutator

(c(ffi), b(a 2™5(cri1 - a > --X -A ccnn |0) \ ,, n

which respect respect the the SL(2, SL(2, C) C) global global conformal conformal invariance. invariance. which Eq. (1.123) (1.123) implies implies the the OPE OPE Eq. c(z)b(w) ~

, (1.125) z— — ww which could also be derived directly from eq. (1.121). In quantum theory, the expression (1.117) for the stress tensor has to be taken in the normally ordered form: Tb,cc(z)=:-Xbdc+(1-X)(db)c: (z) = : -Xbdc + (1 - X)(db)c \){db)c : .


The OPE's between the stress tensor Tbfi and the fields 6, c reproduce the transformation properties of the latter, as they should: dmbH, TbiC(z)b(w) ~ ~ 7 j z b-b(w) ab{w) TUz)b{w) p 6 H++-!—d 7~; '

Tb,c{z)c{w) (z)c(w) ~

(z — w)'


l - A


C ^c(w) «c(w) H N2 2z (™) +

w) (z — w)

(1.127) ( L 1 2 7 )

c(w) . d wc{w) z—ww

Nothing beyond the Wick theorem is needed to check these relations. We are now in a position to calculate the OPE of Tb — 1 have no poles at z = 0 while the unit operator is a scalar, so that one has in|0> Ln |0) = 0 ,, 35

n > -- 1l . .







= 0. The L_„ |0) for n > 2 are the Eq. (2.1) implies a regularity of T(z) |0) at z = non-trivial states. The CFT instates are naturally defined by applying the CFT operators to the vac­ uum, \AiB)=lmioA(z,z)\Q) . (2.2) A(z,2)\Q) To introduce 'out'-states, first we use eq. (2.1) to get (0|L_ n = 0 ,

n > - l .


In CFT's, conformal invariance relates a parametrization near oo on the Riemann sphere C U oo to that of a neighbourhood about the origin via the map z = \/w. Hence, if the A(w, w) are the operators in the coordinates in which w — > 0 corresponds to the point —> at infitiny, it is quite natural to introduce the out-states as follows: 0

We now have to relate A(w,w) to A(z,z). under w — —> 1/wyields yields > z z==1/u;

For primary fields, their transformation law

A{w, w) = A{l/w, l/w)(-w-2)h{-w-2)h



Therefore, the CFT outstates out-states are given by 2h = J jnm^, (01 A(z, z)z2hz2K (Aautl = .


These definitions suggest introducing the notion of the adjoint in CFT, of the form



because of the line of reasoning:

A ( - , - ) —-r—T (Ajutl = limAiw A(w,w)= A(w,w) == lim (0| A


(2.10) (2.10)

where $hj>(z,z) is a primary field. From the OPE between the stress tensor and the primary field $ we get [Ln,n, *$^(z, [L ?)] =f£L = f^L h,K{z, I)]

^T(w)^ w w^T(w)^ hrh(z, h(z,


+1 nn = ([f^^z (z*n+ln -^ = ^ + (n l)z ( zhth,(z,z), z) , = (n +(n+l)z"hj l)zhj hj $ $ aa (z,


so that \h, h) hj = Q, 0, Ln\h,|/i,

n> > 0, 0,

h\h,hj hj = h L00\h,hj \h, h) \h, h) ,

h j -= h \h, h) hj . L0 \h, h)


Eq. (2.12) defines the so-called highest-weight states. The states of the associated highest-weight representation of the Virasoro algebra are created by acting with arbitrary polynomials in {£-„, L-m ; m, nn > > 1} 1} on on the the primary primary state state \h, \h, hj. hj. They They represent represent the the descendant states,

${-%,-%) ss LL__h... i_hki... ${-%,-%) . L__kM h . .L kMl_-

z_^,$ Z_^,$„(0,0) p (o, 0), ,

(2.13) (2.13)

which obviously correspond to the modes of the descendant fields expanded near the origin. A descendant state can be regarded as the result of the action of a descendant field on the vacuum: L_„ \h) = L_„($(0) |0)) = (L_„$)(0) |0) = $M)(wM)) MM n M)(wM)) = £-»tfih)'--fe-iK-i)(iMH)



where the differential operator (n > 2) M-l ' (l-n}hj ^[(l-nfe = £-» = - i=\ £ L[{WJ-IUMY (■Wj WMY i=\ K -—^ M )"

1 d 1 gj' x n 1l {Wj WM) (WJ dwj\ K- -— ■m WM)"dvjj\ u~Y~dWj


has been introduced (n > 2). Eq. (2.18) can be further generalized to the correlation functions of arbitrary descendants. We now want to see how the full conformal invariance in two dimensions restricts the A/-point correlation functions in CFT. As to the 2-, 3- and 4-point functions, they were already considered in sect. 1.2. In particular, eq. (1.45) can be generalized to the A/-point correlation functions by using the conformal Ward identities as follows: n+1 n+1 T(^)0(z)0 (OL_»0(z)fc(zi) S U S ^U (to T( ({L-n)(z)i{z•x)• •■M*M)) ■ ■ 4>M(ZM)) ^ - z)(to - z^TfaWzWi) ■ •■ M)\MZM)\ tU )0(z)01 (z 1 ) • • ■ Mz

M '(l-n)hj 1 =- (_1)»-1£ (-I)"" g [|£g*





(4>(Z)MZI)---4>M(ZM)) "(z(T^RS?] " *" - Zj) ~ dZj_ «')* 7y y == 2/ii. 2—i 2/ij .


In particular, as far as the 2-, 3- and 4-point functions are concerned, the previous results given by eqs. (1.46), (1.47) and (1.49), resp., are reproduced. The OPE coefficients appearing in eq. (2.15) are symmetric, and they coincide with the numerical factors factors in in the 3-point functions of primary or secondary fields, = (#i(oo)4 ($i(oo)$22(*,*)*s(0)) (z,i)§ 3 (0)) ((9i\* * i | 4a(z,S)\* a C M )t)| S 3 ) = hl h2 h3 hi h2 h3

, = C123 z 'l l ~~/ l 2 _z' 1 ''3 ^ 1-"^ 2 _ ^ 3 = 1 o,z'

(2.25) (2.25)

be expressed It is the BPZ theorem [12] that the OPE coefficients for descendants can be for primaries, in terms of those for rP{-i

l i} Cj£*-*> ~ = C^P*pP**> Cmnp/Fj-^^


(no sum over m,nl) m,n\) ,


where the C m n p 's are the OPE coefficients for primary fields, and /?(/?) are some functions of four parameters hm,hn,hp and c (hm,hn,hp and c). The formal proof can be done by performing a conformal transformation of both sides of eq. (2.15) and comparing terms. The /^-coefficients can also be computed this way. Moreover, the 3-point function of any three descendants can be calculated from that of their associated primaries. The primary C m n p 's coefficients actually determine all the 3-point functions for any members of the families [fan], [fa] and [fa], where the [fa] denotes a collection of a primary field and all of its descendants (a conformal family). In CFT, the coefficients Cmnp play the role of Clebsch-Gordan coefficients: given a product of two conformal highest-weight representations, they determine which representations are contained in the decomposition of the product into the conformal irreps [21]. To this end, we outline the derivation of the crucial equation (2.26), following ref. [13]. Unfortunately, there is no explicit formula for a generic coefficient Cmi~k'~h} in this






Fig. 5. For the calculation of the integral on the r.h.s. of eq. (2.27). equation, and the actual calculations become very tedious at high levels (see refs. [12,13] for more). Nevertheless, the general statement represented by eq. (2.26) can be made clear for understanding and practical calulations as well. To simplify the notation even more, the dependence of the kinematical /^-coefficients upon the indices m,n is suppressed. Consider the operator product on the l.h.s. of eq. (2.15b) and apply the operator Ln on it (n > 1): {fa(z)fa{Q)) = j[ c ^C^nOM^MO) Ln(fa(z)fa(0))= ¥-.C+1n$fa(z)fa(0)


Since the operator Ln is defined at the origin and it acts on both fields fa and fa, the contour C should enclose both points z and 0 (see the l.h.s. of Fig. 5). There are now two different ways to proceed with eq. (2.27). The first way is to deform the contour as it is shown in Fig. 5a. Then the -1'-1)2(2A„ 2(2/lp + 1) + ^>3]^5(0) + •. ...} .}, x {0 + z / S ^ / i ^ O ) + ^^[/3,4-^^(0)



^a(^(0)) = pE ^ g ^ {0 + 0 + z2 [p^'-^K + /3 \c)} ^(0) [^-1'-1)6/ip + ( %{Ahp+ J+c )] ^(0) + .+ . . }. ,. . )}


where the definition (1.60) of the descendants, } 4~~k} = $-* = 0 { -""- n2

■ nL"*»> = LL-nimLL-n • • LJpnJp, , n2 2 ■ ■

(2.35) (2.35)

had been used, and the Virasoro operators L\ and L2 have been commuted through those of 0P~*}Comparing eqs. (2.30) and (2.31) with eqs. (2.33) and (2.34), resp., yields the rela­ tions: Pp-l)2hpv = hz2-K -h1 + hp, (2.36a) /3


(4/ip + \c) = = (2/i2 - h /n + hp) . -1



ALGEBRA (2.36c (2.36c)


These equations determine the coefficients f& /3' ',', P Ppp~ '~ /5p~'.'. When comparing ' ^' and pjr coefficients at higher power terms in the above equations, all the /?-coefficents can, in principle, be determined. It is sufficient to consider only the action of the operators L\ and L 2 on a product of two primary fields, since the action of the other Virasoro operators follows, when using the Virasoro algebra. We conclude that the CFT is completely specified by the central charge c, the conforconfor­ mal weights (anomalous dimensions) {h{, hi} of the Virasoro highest-weight states and the OPE coefficients (couplings) Cmnp between the primary fields. The latter are not fixed by conformal invariance, and they have to be determined from dynamics of the underlying physical system. Exercises # II-2 > (*) Consider the conformal Ward identity (1.69) in the limit z — ► — > WM, and derive eq. (2.18) from it. Hint: expand eq. (1.69) in powers of (z — WM), U>M), and use eqs. (1.56) and (1.57). # II-3 t> Check eq. (2.18) directly, by commuting Z_ n to the left on the l.h.s. Hint: take WM = 0 in eq. (2.18) and use eq. (2.11). # II-4 > Take the section z = z" and prove that this leads to the spin quantization, h — h e Z, when requiring the correlation functions be single-valued. Is there a similar restriction on the scaling dimension h + hl


Conformal bootstrap

So far we only considered the local constraints imposed by the conformal algebra. To find the allowed dimensions ft's and couplings Cj^'s in CFT, we need some dynamical principles or some global constraints on the correlation functions. These constraints arise from the associativity of the operator algebra in eq. (1.65), or in eq. (2.15). To see this, we study a 4-point function z (* z4, we find the result schematically pictured on the l.h.s. of Fig. 6, where the sum (over p) goes over both primary and


ti C




iI =:

^ ijpQmpy —p—C 2^CilqCjmq j



/ j

\ m

Fig. 6. The crossing (duality) symmetry of a 4-point function. secondary fields. Second, taking another limit Z\ —*• z3 , z2 —^ —► z^, %i,we wecan canalternatively alternatively evaluate eq. (2.37) with the result diagrammatically represented on the r.h.s. of Fig. 6. The associativity of the operator algebra means that these two methods should give rise to the same result (Fig. 6). This consistency condition is known as the crossing or duality symmetry of the 4-point functions. In QFT, an s-channel Feynman diagram ' has to be supplemented by the similar t- and u-contributions in order to describe a full scattering amplitude. The duality means that the s-contribution alone represents the full amplitude. The relation shown in Fig. 6 places the infinite number of restrictions. They become the algebraic constraints Cyfc's must satisfy, after performing (in principle) the infinite sum over all descendant states. This procedure is known as the conformal bootstrap. In practice, it is very difficult to implement, even in two dimensions where the situation is considerably simpler since we only need to consider primary fields. Fortunately, there are the special values of c's and h's where things are dramatically simplified [12], as we are going to discuss in what follows. To convert the diagrammatic relation of Fig. 6 to an analytic expression, we choose three reference points to be zx = = 00, oo, z2 = 1 and 2:4 = 0, while z3 = = z. We are now interested in the Green's function fc \Z) **>-*»-*«« -k" ^ lp)-Ifl

Ji A* Ji

1 - z. Similarly, when z2 -> under the SL(2, SL(2, C) transformation -> 1/z, 1/z, one has the transformation law 21 G^{Z,Z): = ^z-^z-^GZl ' ^ gQ ,, j ) .■ (2.45) G lm((z, z , i) z) = G? "z'^z-' ^ " G

•cS(i.i) ■

A graphical representation of eq. (2.43) is given in Fig. 8. This equation is apparently first to be computed which is a very complicated: the conformal blocks themselves are first non-trivial task, while the sum goes over all the 3-point couplings which are infinite in number. The conformal blocks are determined by the conformal Ward identities. Given all the conformal blocks J-', !F, eq. (2.43) places the set of very non-trivial equations characterizing all possible OPE coefficients. The problem of classifying all 2d CFT's is therefore equivalent to solving eq. (2.43) for all possible values of the central charge c (and adding the modular invariance condition, see Ch. III). At the special values of c and h corresponding to the minimal models, the conformal blocks can be determined as solutions to the linear differential equations resulting from a presence of certain null states (see next sections). An investigation of general corollaries of eq. (2.43) is one of the central problems in CFT. The two basic properties of CFT we have used so far were (i) locality, and (ii) as­ sociativity of OPE's. Although conformal blocks are generically multi-valued, physical correlators must always be single-valued. The multi-valuedness or the non-trivial monodromy of conformal blocks means that they are not just ordinary functions, but rather sections of some non-trivial vector bundle. The notion of duality can be generalized to the M-point correlation functions (and, in fact, to arbitrary-genus Riemann surfaces as well). A general correlation function Giy..iM {Z\,...,ZM) can be expressed in terms of more general conformal blocks pictured in Fig. 9.







'M-2 1M-I

Fig. 9. The graphical representation of a general conformal block. Any physical correlator has to be a monodromy-invariant combination of Tpx and Tfc, of the form Gi,...lM(zuz1;...;zM,zM)

= £


(zu .. .)J%"i" {zx,...)




where an invariant metric hfltp2 is to be constructed out of the structure constants (cou­ plings) Cijk's. Quite generally, the Green's function G should be (i) a single-valued real analytic function of the coordinates zt's and moduli of the Riemann surface, (ii) inde­ pendent of the order of its arguments, and (iii) independent of the basis of conformal blocks used to compute it. As far as conformal blocks are concerned, this is not true: a function T is, in general, multi-valued. It is the monodromy (or braid) transforma­ tion laws, that are to be used to define T for some ordering of its arguments, and to analytically continue it to different domains (sect. XII.6).


Null states

'Solving' a CFT means finding its full spectrum and all correlation functions. Having the infinite-dimensional conformal symmetry, the general strategy towards formulating solvable CFT's is to impose as many constraints as possible, in a way to be consistent with this symmetry. This way to proceed is just one of the ways to realize the 'boot­ strap' idea: imposing all possible conformal constraints should result in not just very special CFT's but, presumably, the all of them which are internally consistent, with no additional symmetries used. The closely related idea is to start to look for unitary representations of the Virasoro algebra. Each CFT state can be expressed as a linear conbination of primary and secondary states. The only relevant representations having a chance to contain unitary Virasoro irreps are just the highest-weight representations, since they have 'energy' levels


49 3

bounded from below. Starting from a highest-weight state \h), we can build up the associated Verma module. However, this module does not necessarily have a positively definite scalar product, i.e. it may not be a Hilbert space, in general. This depends on the structure of the Verma module for given values of c and h. The descendant state |x) satisfying the equations L0 Ix) = (h + N) |x) \X) ,

Ln |x) \X) = 0 for n > 0 ,


is known as a null state or a singular vector. The null state is simultaneously a primary and a descendant state, and it is a highest-weight state also. To obtain a nondegenerate representation, one has to eliminate all null states and all of their descendants, and consider the reduced theory. Later we will see how it helps to evaluate conformal blocks, which form a (fully reducible) representation of the duality transformations. The elimination of null states will result in non-trivial linear differential equations for conformal blocks. We are now looking for degenerate representations of the Virasoro algebra which contain null states. Eq. (2.47) can be equally interpreted as the requirement of conformal invariance of the constraint |x) = 0 we want to impose. Consider a Verma module V(c, h) of the Virasoro algebra, with a highest-weight vector \h). If there is a null vector |x) satisfying eq. (2.47), this means that {ip | x) = 0 for any state (4>\ of the theory. Setting |x) = 0 is equivalent to taking the quotient by the null vectors and submodules they generate. At the level one, the only possibility is to take |x) = i - i \h) = 0 or L-ih(z) = dzh{z) = 0 (see the third line of eq. (1.59)). This obviously leads to the identity operator and \h) = |0), which is a trivial representation. At the level two, it may happen that \X) = L-2\h)+aL2_l\h)=0


for some value of the parameter a. At least, eq. (2.48) respects the scaling symmetry because both terms on the l.h.s. have the same scaling dimension. Since all Ln operators of the Virasoro algebra at n > 0 are generated (in the algebraic sense) by L\ and L2, applying L\ and L2 to eq. (2.48) is enough to ensure the whole conditions for a null state in eq. (2.47). 3

The 'energy' levels are always the eigenvalues of the Hamiltonian. In the case under consideration, they coincide with the 'levels' - the eigenvalues of the operator L0.






Acting with L\ 2a(2h + Lx on eq. (2.48) and using usirig the Virasoro algebra yields [3 + 2a(2n l)]L_i \h) = 0, whereas acting with L22 res results in [Ah + c/2 + 6ah] \h) = 0. Hence, we ults [4/i + c/2 + 6ah] \h) = 0. Hence, we find 2 / l 3 ,, 4cn ^ - -8 8/t \ (2.49) c = 2/i(5 (2 49)' 2(2/i+l) ' C = ^ 2^/ i + ° = -2(2ftTI)' Tl ^ ' ' and, therefore, therefore, and,


^V-K i^-'-^W-'K \XK)=(L-2

where where



A = - ^ ( 5 - c ± > //((cc-- ll ) ( c - 2 5 ) ))

(2.51) (2.51)

Eq. (2.50) can be rewritten to the form L Mz) = L_ L-22 Xh{z) = h{z)~- 2(2h+l) 2(2*+l)L-iM») jft(«) 0„(z) -'


=° 0• •

( 2(2.52) 52)


We are now going to show that eq. (2.52) leads to a second-order partial differential field K(Z). equation for the correlation functions containing the degenerate primaryfield h.{z). Eq. (2.52) implies that the correlation functions of h with other fields are annihilated by the differential operator £_ 2 — 2(2fe+i)^-i D e c a u s e of eq. (2.18). Since

L-i&M = §;k]'s denote primary fields, non-negative integers, which to0fc. interpreted as the numbers of independent fusion paths from 0, and 0 ; to fa. Clearly, Nijk = 0 implies C*/ 0. The inverse statement is also true [52]. This property of C„* = 0. CFT is known as naturality.

The coefficients JVy* Mi/ are automatically symmetric in i and j , and satisfy a quadratic equation due to associativity of the multiplication in eq. (2.77). It can be shown [53]



that for CFT with a finite number of primaries (i.e. for a RCFT), all the dimensions hi are rational numbers. The minimal models with 0 < c < 1 are the examples of RCFT. Another important class of RCFT's is provided by the WZNW theories (Ch. IV). The rationality condition means the indices of JVy* are running over a finite set of values, and, hence, summations over them are well-defined. Using the matrix notation, (Ni)jk = iVy*, the associativity of the OPE implies [Ni,Nj\ = 0 or NtNj = T.kNijkNk- Therefore, the matrices N JVJ themselves themselves form form aa commutative commutative and and associative associative matrix matrix representation representation of of the the fusion rules (2.77). In a RCFT, the Virasoro algebra may be extended to a more general symmetry algebra, normally represented by a W algebra (Ch. VII) and called a chiral algebra (see Chs. IV and VII for more). The problem of classifying all RCFT's is therefore reduced to finding all possible chiral algebras and automorphisms of the associated fusion rules [52]. Unfortunately, there are no general rules to determine the fusion rules in CFT.


Coulomb gas picture

As was shown by Dotsenko and Fateev [54, 55], all the essential features of the minimal models can be obtained by using a single scalar field interacting with a background charge. This approach is known as the 'Coulomb gas' or Dotsenko-Fateev (DF) construction. Their method is very effective in computing correlations functions, as will be illustrated on a particular example in sect. 7. Let us consider a reparametrization-invariant 2d action of a bosonic scalar field $(2, -foo), $(z, z) on the Riemann sphere S 2 ~ C U {oo}, SSDF($) DF($) = g^ / (fz ** ((9z*3 lioL^m) = ^J 4 * *z-$ * ++ 2ic* ov^R$) ,,


where the additional term proportional to the 2d (scalar) curvature R has been intro­ duced. In eq. (2.78), a 0 is a real constant. The correlation functions are formally defined by eq. (1.3) with the action (2.78). The curvature of the sphere can be concentrated at the north pole, representing the point at infinity, by the appropriate choice of metric. The Riemann sphere can then be described by the two coordinate charts, the nothern one in the vicinity of the north pole and the southern one everywhere outside it. For the latter, we can take the metric in the form (1.29), just as that for a plane. The curvature term in the action (2.78) results






in the additional term in the stress tensor having the form 222 TDF(Z) = = - \ : dip(z)dcp(z) dtp(z)d 1 , a = l, (2.111) with the associated dimensions with the associated dimensions h(n,m) = a2nm - 2a n , m a 0 h(n,m) = a2nm - 2a„ jm a 0 (na + + 77ia_)22 + (a + — a_) 22 _ (na + + 77ia_) + (a + — a_) 4 ~ 4 = - aa 2 + \{na+ + ma-)22 , (2.112) = ~ l + \{na+ + ma-) , (2.112) once once again again reproducing reproducing the the Kac Kac spectrum spectrum (2.59) (2.59) of of the the degenerate degenerate conformal conformal families! families! Given Given eq. eq. (2.112), (2.112), the the 4-point 4-point correlation correlation functions functions are are given given by by analytic analytic integrals integrals in in the the so-called so-called Feigin-Fuchs Feigin-Fuchs representation: representation: *^


i> du\--- dum-idum-ii

du\ ■ ■ ■ JCi


JC Jcmm-i-i



dvi--- dv\

■/s„_ 7s„_iI

dvdv n-i n-i

■ ■J-{u ■ m-i)J+{vi)--



An explicit calculation of the correlator can be completed by using integral representa­ tions of the hypergeometric functions [54] (see sect. 7 for an example). As an application of this construction, let us derive the fusion rules (2.76). To find the fields 0(fci() appearing in the OPE for the product of 4>(m,n) and 0(r,3), consider the 3-point function in the Coulomb gas picture, and use SL(2, C) invariance to write down three equivalent ways to represent it: V (Vfeo(oo)V < (M)(°°),l/(m,n)(l)V( r,(0)Q+Qi-) 3)(0)Q+g:-) , (ra , B) (l)V M


((W~)fe(i)Wo)Q+Q-) W ~ ) f e ( l ) W 0 ) Q + Q - ) -, ((v VM ( o o ) VM ( l ) ^R ( 0 ) ( 9 ^ : ) ,, M (oo)v M (i)^ R (o)(9^:)

(2.1146) (2.114c)

11.6. PICTURE II.6. COULOMB GAS PICTURE where V(m,n) = V a(m , n) and V^tf Vj^ the correlator will not vanish if


= =V ^ , , - ^(m-nV V2aa-a (2.111) and and (2.114a), n) . According to eqs. (2.111)

k < m + r — 11 with k — m — r — 1 even, (2.115)

I ^1,203,1 > ^2,201,1 + ^2,203,1 where C\2 = 1, as it should. What is important is that the C\2 vanishes when 7 = 4/3, because of the pole in one of the T's appearing in eq. (2.152a) for the a3\ = a3 . Indeed, a3 then becomes divergent and, hence, C 3^^ vanishes due to eq. (2.151). But this just means that the operator 0 3il decouples. Similar things happen in the other minimal models (although, in a less obvious way as above) [13, 65]. The minimal models are in fact the only consistent CFT's which can be constructed this way by the conformal bootstrap! To see this, let us take a look at the Kac formula (2.112) again, and notice that some of the values of h(n,m) become negative near the line in the (n, m) plane where a+n + a^m a..m = a+n — —|a_|n |a_|n==00. .

(2.157) (2.157)

The 2-point functions of the corresponding primary operators grow with distance and, hence, they are to be considered as 'pathological' in the sense of the ordinary field n

F o r simplicity, the scaling factors are dropped in what follows, see eq. (1.65).






theory. Therefore, it makes sense to get rid of such operators in CFT. Having a set of 'good' primaries with positive dimensions, their mutual OPE's could generate some new operators which have to be included into the set in order to close the operator algebra. Unless the operator algebra closes on a finite set, among the new operators to be ultimately included there will be some with negative dimensions, thus making the theory inconsistent. In particular, only the family with q = p + 1 has all positive conformal dimensions (exercise 11-14). All the other models fail to be unitary [48]. This explains why we have chosen to stick to the RCFT's or just to the minimal models, since they seem to be the only ones which are fully consistent as QFT's. The unitarity of CFT may not be so crucial in statistical physics, where correlation functions may grow with distance. Exercises # 11-13 > (*) Consider the spectrum of the tricritical Ising model from the exercise 14: operators. It has the generic form 11-11, and calculate the operator product of two 0cf> 3il 03,13,1 ^3,303,1 + £>3,305,1 3,13,1 ~ £>3,3^U £>3,33>3,1 -C»3,3^5,l ■•

(2.158) (2.158)

Verify that D\ - coefficient in 2?|33 is finite, but D\ D$3 isis vanishing. vanishing. Hint: consider the a5-coefEcient eq. (2.152a) and prove that it diverges at 7 = 6/5. # 11-14 > (*) Consider the rational values of 7 =- a\ = q/p and prove that conformal dimensions inside a finite (q — 1) X (p — 1) tableare ;areallallpositive positivewhen whenq q= =p p+ +1,1,but but have negative values when q / p + + 1. Hint: use the Kac formula (2.112).


Felder's (BRST) approach to minimal models

The rhe Fock space of states of a minimal model is the direct sum of some finite number of irreducible rreducible highest weight modules of the Virasoro algebra Vir © Vir, 12 U — ©,«„,„. ®,rLnni ® ® tinn* W= 12



0ur Our notation in this section is very close to that used in Felder's original paper [66], [66], although Our some efforts have been made to make it compatible with ours ours.:. In particular, Vir = {Lk} {Lk} = {{LLk t ® l1} } and Vir = {Lk} = {1 ® Lk}. ). The two pairs of integers used to parametrize the central charge c and the highest weight spectrum {/i„,„ ip,P') and (n,m) ->





Each module HHnn,,n n' in eq. (2.159) is degenerate in the sense that it is given by a quotient of the Verma module V(hn«]

+ K A k

+ 1)zk


^« © HH- -Vv t^f )> M W

( 2 - 161 )

The constants Ckjum>mn represent finitely many 3-point function coefficients (CFT structure constants). The normalization of the block fields \nn'\(m,m'\( \nn> ••

2 18 ((2.180) - °)

where I = n + m — IT — 1 and /' = n' + m' — 2r' — 1. It is straightforward to compute eq. (2.180) by reducing it to one of the standard DF integrals [55], like those considered 16

The phase in the first equation (2.178) appears as the result of commuting K«.,,i with all the Ka+ 's comprising the BRST charge Qm.






in sect. 7. The value of this constant turns out to be non-vanishing only if the minimalmodels fusion rules (2.119) are satisfied [66] ! Simultaneously, eqs. (2.163) and (2.172) provide the integral representation for the conformal blocks, * % j ( * i , . . -. , * ) = const < v*u, [ ] t £ ' 2 ( a , ) » n > , ?&{zu


where kj~\ = n;- -1- m, - 2r7- — 1 and fcj_i = = n'j + m'j — 2rJ — 1. The contours drawn in Fig. 13 can be deformed in such a way that the conformal blocks (2.181) become linear conbinations of the DF-type integrals (sect. 7). This completes the Felder's proof of the Feigin-Fuchs integral representation on a plane. In the case of the WZNW theories (Ch. IV), this representation is proved in sect. IV.4. Felder's constriction of the minimal models can be straightforwardly generalized to Riemann surfaces of arbitrary genus. The case of torus (genus-1) will be considered in sect. III.5. Since Felder's approach is based on the fundamental BRST structure [67, 68], one may ask whether his construction originates from the gauge fixing of an underlying gauge theory. This gauge theory, if any, is yet to be found. To conclude this section, we note that although the minimal models are the sim­ plest examples of CFT indeed, from the viewpoint of their BRST structure they are simultaneously the most complicated ones because of their highest level of degeneracy. Using free field realizations is very useful in analyzing CFT's, and it always implies three necessary ingredients [69]: • realization of CFT chiral operator algebra A by free fields, • 'null-state decoupling' by projecting the free field Fock spaces to the irreps of A, • realization of the vertex operators. The BRST structure should also exist in any CFT formulated on an arbitrary Riemann surface, with the BRST symmetry playing a fundamental role.

Chapter III Partition Functions and Bosonization In this Chapter we consider CFT's on Riemann surfaces of genus h > 1. The case of torus, h = 1, will be most important for us. A consistent formulation of CFT on a torus, or on a Riemann surface of arbitrary genus, normally gives rise to further constraints on CFT operator contents. The string perturbation theory (Ch. VIII) gives the natural area for applications. Chiral bosonization on Riemann surfaces and Felder's (BRST) formulation of the minimal models on a torus are also considered in this Chapter.


Free fermions on a torus

The general strategy for constructing CFT's on a torus is to make use of the local properties of the CFT operators already constructed on a conformal plane, map this plane to a cylinder via the exponential map, and then make the torus via discrete identification. Though this procedure preserves all local properties of the operators, it does not necessarily preserve all of their global symmetries. A torus maps to an annulus on a plane, so that only LQ and Lo survive as the global symmetry generators. The global symmetry group is therefore reduced to t/(l) x t/(l). Boundary conditions on conformal fields are also affected by the passage from the plane to the cylinder (or to the torus). A primary field 4>hjXz,z) transforms under the 83






—> z = = e as map w —>


M) = M)


1 (3.1) (3)

A*K «) = ( ^ ) (;§) ^- *) = ^ C * . * )

0cyl(™, w) =

*'z, same Hence, under 2z —> e22,rl z, CFT primary fields with integer spin s = h — h have the some boundary conditions on the plane and the cylinder. On the contrary, CFT primary fields with half-integer spin change the periodic boundary conditions to the anti-periodic ones and vice-versa, when passing from the plane to the cylinder, as we have already noticed in eqs. (1.104) and (1.105). As far as the CFT stress tensor T(z) is concerned, which is not even a tensor under the conformal transformations, the anomalous piece proportional to the Schwartzian derivative arises: w w w^z ->z = = eTO : S(ew,w) = -l/2

TcyiH == TcylH


T{z)+ s{z w)=z2T{z)= z2T{z)

+ (£) T2'S{z'w) (£) T{z) T2

~~2I k•

(3 2) (3.2)


Substituting the mode expansion T{z) = J2n Lnz~n~2 into eq. (3.2) yields Tcyl(w) = £ Lnz~n - ± = £ ( n€Z

L n

- y46n,o)e-™ ,





and, and, hence, hence, (3.4) (3.4) The shift of the vacuum energy does make definite sense in CFT since the scale and rotational invariances on the plane naturally fix the La and L0 eigenvalues to be zero for an SL{2, C)-invariant vacuum. (io)cyl = L L0 -— iL — .. (L ) = 0 cyl


To introduce the CFT partition function on a torus, let us define first the real Hamiltonian (H) and momentum (P) operators as the generators of translations in the 'time' and 'space' directions, resp. It is convenient to redefine at this point the coordinate variable, w — —>> iw, so that Re(iu) and Im(tu) now become the space and time coordinates, ^0) generate dilatations and rotations on the plane, one can identify H resp. Since (Lo ± £0) (^o)cyi, and P with the imaginary part of (L (Lo)cyi (^o)cyi- For the torus with (Lo)cyi + (-t'o)cyi, 0)cyi - (io)cyiwith a modular parameter r = rj + ir2, we identify two periods of w as w = w + 2tr and w = w + 2ITT (or, equivalently, Im(ui) = lm(w) + 2irr2 and Re(u>) = Re(ui) + 2TXT{). Therefore, the operator which generates translations in the Euclidean (imaginary) time is e~2lTT2H, whereas the accompanying translation operator in the Euclidean space is e-27rr1p -phis suggests a definition of the partition function in the form: 2 T2H e --2-irriP Z(T)=Tre* -2nre-2H2"TiP Z(r) = Tre-




cyl-'^°' 0 cy1 cy1' Tre 2lr2lr* TlTl" LoL ' cyl_

= TVe * " °' = 2ltiT L

'^ '

_22wr rT2 Lo L C) 1+ 0 c) 1

" o)cy\H^)cy\] ' ' ^' 'l ' ee~' l[(

-_22 r i FF



' ' ((£o)eyl ^ o ) c y l _ >JY„(£o)cylg(£o)cyl = 'YJTl eg2irtT(Lo) ^ °^ g '™' J> „( o)cylg 0) to the highest weight state. Since [Lo,I/_„] = nL_ n , the character of V is

v(9) v(9)=== xv(9)


where ]T£Li(l

ft-c/24 qh-c/U

fiE^wr nE^wr n^d-f)'

(3 9) (3 9) (3.9)


-1 n an