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This book describes the practical techniques for connecting the phenomenology of particle physics with the accepted modern theory known as the 'Standard Model'. The Standard Model of elementary particle interactions is the outstanding achievement of the past forty years of experimental and theoretical activity in particle physics. This book gives a detailed account of the Standard Model, focussing on the techniques by which the model can produce information about real observed phenomena. The text opens with a pedagogic account of the theory of the Standard Model. Introductions to the essential calculation techniques needed, including effective lagrangian techniques and path integral methods, are included. The major part of the text is concerned with the use of the Standard Model in the calculation of physical properties of particles. Rigorous and reliable methods (radiative corrections and nonperturbative techniques based on symmetries and anomalies) are emphasized, but other useful models (such as the quark and Skyrme models) are also described. The strong and electroweak interactions are not treated as independent threads, but rather are woven together into a unified phenomenological fabric. Many exercises and diagrams are included.
CAMBRIDGE MONOGRAPHS ON PARTICLE PHYSICS, NUCLEAR PHYSICS AND COSMOLOGY 2 General Editors: T. Ericson, P. V. Landshoff
DYNAMICS OF THE STANDARD MODEL
CAMBRIDGE MONOGRAPHS ON PARTICLE PHYSICS, NUCLEAR PHYSICS AND COSMOLOGY 1. K. Winter (ed.): Neutrino Physics 2. J. F. Donoghue, E. Golowich and B. R. Holstein: Dynamics of the Standard Model 3. E. Leader and E. Predazzi: An Introduction to Gauge Theories and Modern Particle Physics, Volume 1: Electroweak Interactions, the "New Particles" and the Parton Model 4. E. Leader and E. Predazzi: An Introduction to Gauge Theories and Modern Particle Physics, Volume 2: CP Violation, QCD and Hard Processes 5. C. Grupen: Particle Detectors 6. H. Grosse and A. Martin: Particle Physics and the Schrodinger Equation
DYNAMICS OF THE STANDARD MODEL
JOHN F. DONOGHUE EUGENE GOLOWICH BARRY R. HOLSTEIN University of Massachusetts, Amherst
I CAMBRIDGE UNIVERSITY PRESS
Published by the Press Syndicate of the University of Cambridge The Pitt Building, Trumpington Street, Cambridge CB2 1RP 40 West 20th Street, New York NY 100114211, USA 10 Stamford Road, Oakleigh, Melbourne 3166, Australia © Cambridge University Press 1992 First published 1992 First paperback edition (with corrections) 1994 Reprinted 1996 A catalogue record for this book is available from the British Library Library of Congress cataloguing in publication data available ISBN 0 521 36288 1 hardback ISBN 0 521 47652 6 paperback
Transferred to digital printing 2002
UPH
To our families
Contents
I 1.1 1.2
1.3
1.4
1.5
II II. 1
11.2
11.3
Preface Inputs to the Standard Model Quarks and leptons Chiral fermions The massless limit Parity, time reversal, and charge conjugation Symmetries and near symmetries Noether currents Examples of Noether currents Approximate symmetry Gauge symmetry Abelian case Nonabelian case Mixed case On the fate of symmetries Hidden symmetry Spontaneous symmetry breaking in the sigma model Interactions of the Standard Model Quantum Electrodynamics U(l) gauge symmetry QED to one loop Onshell renormalization of the electric charge Electric charge as a running coupling constant Quantum Chromodynamics SU(3) gauge symmetry QCD to one loop Asymptotic freedom and renormalization group Electroweak interactions Weak isospin and weak hypercharge assignments
ix
xvii 1 1 4 4 6 8 8 9 13 14 14 15 17 18 19 21 24 24 25 27 32 33 36 36 41 47 52 52
x
Contents gaugeinvariant lagrangian Spontaneous symmetry breaking Electroweak currents Fermion mixing Diagonalization of mass matrices Quark mixing CPviolation and rephasinginvariants SU(2)LXU(1)Y
II.4
III Symmetries and anomalies 111.1 Symmetries of the Standard Model 111.2 Path integrals and symmetries The generating functional Noether's theorem and path integrals 111.3 The U(l) axial anomaly Diagrammatic analysis Path integral analysis 111.4 Classical scale invariance and the trace anomaly 111.5 Chiral anomalies and vacuum structure The 0vacuum The 0term Connection with chiral rotations
54 56 58 60 60 63 65 69 69 72 73 74 75 77 81 88 91 92 94 95
IV Introduction to effective lagrangians IV. 1 Nonlinear lagrangians and the sigma model Representations of the sigma model Representation independence IV.2 Integrating out heavy fields The decoupling theorem Integrating out heavy fields at tree level IV.3 The low energy expansion Expansion in energy Loops Weinberg's power counting theorem IV.4 Symmetry breaking IV.5 PCAC The softpion theorem IV.6 Matrix elements of currents Matrix elements and the effective action IV.7 Heavy particles in effective lagrangians IV.8 Effective lagrangians in QED IV.9 Effective lagrangians as probes of new physics
97 97 98 100 102 102 103 105 105 106 108 109 111 111 113 114 116 119 121
V V.I
126 126
Leptons The electron
Contents
xi
BreitFermi interaction QED corrections The infrared problem The muon Muon decay at treelevel Photon radiative corrections The tau Inclusive decays Exclusive leptonic decays Exclusive semileptonic decays The neutrinos Neutrino oscillations Terrestial searches for neutrino mixing Solar neutrinos Dirac mass and Major ana mass
126 129 134 136 137 138 143 144 145 146 149 150 151 151 154
VI Very low energy QCD  pions and photons VI. 1 QCD at low energies Vacuum expectation values and masses Pion leptonic decay and Fn VI.2 Chiral perturbation theory to one loop The order E4 lagrangian The renormalization program VI.3 Interactions of pions and photons The pion form factor Rare pion processes VI.4 Pionpion scattering VI.5 The axial anomaly and TT°—> 77 VI.6 The physics behind the QCD chiral lagrangian
157 157 158 159 163 164 166 170 170 174 177 181 183
VII Introducing kaons and etas VII. 1 Quark masses VII.2 Higher order analysis of decay constants and masses Ambiguities in mass parameters Decay constants Masses VII.3 The WessZuminoWitten anomaly action VII.4 The T/(960) rjo  7/8 mixing
188 188 191 192 193 194 196 201 204
VIII Kaons and the AS = 1 interaction
208
VIII. 1 Leptonic and semileptonic processes Leptonic decay Kaon beta decay and V us
208 208 209
V.2
V.3
V.4
xii
Contents
The decay K > 7r7rePe VIII.2 The nonleptonic weak interaction VIII.3 Short distance behavior Short distance operator basis Perturbative analysis Renormalization group analysis VIII.4 The A/ = 1/2 rule Phenomenology Chiral lagrangian analysis Vacuum saturation VIII.5 Rare kaon decays
212 213 216 216 217 218 222 222 225 227 228
IX Kaon mixing and CP violation IX. 1 K°K° mixing Mass matrix phenomenology Box diagram contribution IX.2 The phenomenology of kaon CP violation IX.3 Kaon CP violation in the Standard Model Analysis of e Penguin contribution to e1 Additional contributions to ef IX.4 Electric dipole moments IX.5 The strong CP problem The parameter 0 Connections with the neutron electric dipole moment
232 232 232 235 238 242 243 243 246 250 252 253 254
The TV"1 expansion The nature of the large Nc limit Spectroscopy in the large Nc limit Goldstone bosons and the axial anomaly The OZI rule Chiral lagrangians Weak nonleptonic decays
258 258 260 263 265 267 269
XI Phenomenological models XI. 1 Quantum numbers of QQ and Qs states Hadronic flavorspin state vectors Quark spatial wavefunctions Interpolating fields XI.2 Potential model Basic ingredients Mesons Baryons Color dependence of the interquark potential
273 273 274 277 279 280 281 282 283 284
X X.I X.2 X.3 X.4 X.5 X.6
Contents
xiii
XI.3 Bag model Static cavity Spherical cavity approximation Gluons in a bag The quarkgluon interaction A sample fit XI.4 Skyrme model SineGordon soliton Chiral 51/(2) soliton The Skyrme soliton Quantization and wavefunctions XI.5 QCD sum rules Correlators Operator product expansion Master equation Examples
285 286 286 289 290 291 292 292 293 295 297 302 302 305 306 307
XII Baryon properties XII. 1 Matrix element computations Flavor and spin matrix elements Overlaps of spatial wavefunctions Connection to momentum eigenstates Calculations in the Skyrme model XII.2 Electroweak matrix elements Magnetic moments Semileptonic matrix elements XII.3 Symmetry properties and masses Effective lagrangian for baryons Baryon mass splittings and quark masses GoldbergerTreiman relation The nucleon sigma term Strangeness in the nucleon Quarks and their spins in baryons XII.4 Nuclear weak processes Measurement of Vud The pseudoscalar axial form factor XII.5 Hyperon semileptonic decay XII.6 Nonleptonic decay Phenomenology Lowestorder chiral analysis Quark model predictions
313 313 313 315 316 319 321 321 324 325 326 327 329 330 332 333 336 336 337 339 341 341 343 345
XIII Hadron spectroscopy XIII.l The charmonium and bottomonium systems
350 350
xiv
Contents
Transitions in quarkonium XIII.2 Light mesons and baryons SU(6) classification of the light hadrons Regge trajectories SU(6) breaking effects XIII.3 The heavyquark limit Heavyflavored hadrons in the quark model Spectroscopy in the TRQ —• oo limit XIII.4 Nonconventional hadron states Gluonia Additional nonconventional states
356 360 360 363 365 370 370 372 375 376 379
XIV Weak interactions of heavy quarks XIV.l Heavy quark lifetimes and semileptonic decays The spectator model Beyond the spectator model Inclusive vs. exclusive models for b —• ceue The b —• ueve transition The top quark XIV.2 Weak decays in the heavyquark limit XIV.3 B°B° and D°D° mixing B°B° mixing D°D° mixing XIV.4 The unitarity triangle XIV. 5 CP violation in B mesons CPodd signals induced by mixing CPodd signals not induced by mixing XIV.6 Rare decays of B mesons The quark transition b —• 57 The hadron transition B > K*y
382 382 382 384 386 387 389 391 395 395 398 399 400 401 407 409 410 412
XV The Higgs boson XV. 1 Mass and couplings of the Higgs boson The naturalness problem XV.2 Production and decay of the Higgs boson Decay Production XV.3 The possibility of a strongly interacting Higgs sector The equivalence theorem Scattering of longitudinal gauge bosons
415 415 418 419 419 420 423 424 427
XVI The electroweak gauge bosons XVI. 1 Neutral weak currents at low energy Neutral current effective lagrangians
430 430 431
Contents
xv
Determination of the weak mixing angle 0W Definitions of the weak mixing angle Phenomenology of the W± and Z° gauge bosons Decays of W± into fermions Decays of Z° into fermions Asymmetry measurements The number of light neutrino generations The VWW vertex The quantum electroweak lagrangian Gaugefixing and ghost fields in the electroweak sector A subset of electroweak Feynman rules Onshell determination of electroweak parameters Selfenergies of the massive gauge bosons The charged gauge bosons W^ The neutral gauge bosons Z°, 7 Examples of electrowweak radiative corrections The O{G^ml) contribution to Ap The O(GMra?) contribution to Ar The Z —> bb vertex correction Beyond the Standard Model
432 435 439 440 442 442 445 446 448 448 449 451 452 453 455 456 457 458 460 461
Appendix — Functional integration Quantum mechanical formalism Path integral propagator External sources The generating functional The harmonic oscillator Field theoretic formalism Path integrals with fields Generating functional with fields Quadratic forms Background field method to one loop Fermion field theory Gauge theories Gauge fixing Ghost fields
468 468 468 471 472 474 476 477 478 480 482 483 486 487 490
B B.I B.2
Appendix — Some field theoretic methods The heat kernel Chiral renormalization and background
495 495 499
C C.I C.2
Appendix — Useful formulae Numerics Notations and identities
XVI.2
XVI.3
XVI.4
XVI.5
XVI.6 A A.I
A.2 A.3
A.4 A.5 A.6
fields
505 505 505
xvi C.3 C.4 C.5
Contents Decay lifetimes and cross sections Field dimension Mathematics in ^dimensions
508 511 511
References Index
514 532
Preface
The Standard Model lagrangian £SM embodies our knowledge of the strong and electroweak interactions. It contains as fundamental degrees of freedom the spin onehalf quarks and leptons, the spin one gauge bosons, and the spin zero Higgs fields. Symmetry plays the central role in determining its dynamical structure. The lagrangian exhibits invariance under 577(3) gauge transformations for the strong interactions and under £17(2) x 17(1) gauge transformations for the electroweak interactions. Despite the presence of (all too) many input parameters, it is a mathematical construction of considerable predictive power. There are several books available which describe in detail the construction of £SM &nd its quantization, and which deal with aspects of symmetry breaking. We felt the need for a book describing the next steps, how £SM is connected to the observable physics of the real world. There are a considerable variety of techniques, of differing rigor, which are used by particle physicists to accomplish this. We present here those which have become indispensable tools. In addition, we attempt to convey the insights and 'conventional wisdom' which have been developed throughout the field. This book can only be an introduction to the riches contained in the subject, hopefully providing a foundation and a motivation for further exploration by its readers. In writing the book, we have become all too painfully aware that each topic, indeed each specific reaction, has an extensive literature and phenomenology, and that there is a limitation to the depth that can be presented compactly. We emphasize applications, not fundamentals, of quantum field theory. Proofs of formal topics like renormalizability or the quantization of gauge fields are left to books such as Bjorken and Drell, Cheng and Li, Itzykson and Zuber, Pokorski, and Ramond. We include no analysis of parton phenomenology, and refer the reader, for example, to the books by Barger and Phillips or Field. In addition, the study by computer of lattice field theory is an extensive and rapidly changing disxvii
xviii
Preface
cipline which we do not attempt to cover. Although it would be tempting to discuss some of the many stimulating ideas, among them supersymmetry, grand unification and string theory, which attempt to describe physics beyond the Standard Model, limitations of space prevent us from doing so. Although this book begins gently, we do assume that the reader already has some familiarity with quantum field theory. As an aid to those who lack familiarity with path integral methods, we include a presentation, in Appendix A, which treats this subject in an introductory manner. In addition, we assume a knowledge of the basic phenomenology of particle physics, say, at the level of the books by Perkins or Halzen and Martin. We have constructed the material to be of use to a wide spectrum of readers who are involved with the physics of elementary particles. Certainly it contains material of interest to both theorist and experimentalist alike. Given the trend to incorporate the Standard Model in the study of nuclei, we expect the book to be of use to the nuclear physics community as well. Even the student being trained in the mathematics of string theory would be well advised to learn the role that sigma models play in particle theory. This is a good place to stress some conventions employed in this book. Chapters are identified with roman numerals. In crossreferencing equations, we include the chapter number if the referenced equation is in a chapter different from the point of citation. The Minkowski metric is g^iv = diag {1,1,—1,1}. Throughout, we use the natural units h = c = 1, and choose e > 0 so that the electron has electric charge — e. We employ rationalized HeavisideLorentz units, and the fine structure constant is related to the charge via a = e2/47r. The coupling constants for the SU(3)C x SU(2)L X U(l) gauge structure of the Standard Model are denoted respectively as 53, g(x), C A»(x) C~ l = A"(x) , T A^fx) T" 1 — A (xrp)
P A*(x) P'1 = AJxp) ,
Beginning with the discussion of Noether's theorem in the next section, we shall explore the topic of invariance throughout much of this book. It suffices to note here that the Standard Model, being a theory whose dynamical content is expressed in terms of hermitian, Lorentzinvariant lagrange densities of local quantum fields, is guaranteed to be invariant under the combined operation CPT. Interestingly, however, these discrete transformations are individually symmetry operations only of the strong and electromagnetic interactions, but not of the full electroweak sector. We see already the possibility for such behavior in the occurrence Table 15. Response of Dirac bilinears to discrete mappings S(x) P{x)
S(xP) P{xP)
S(xT) P(xT)
8
/ Inputs to the Standard Model
of chiral fermions II)L,R, since parity maps the fields ^L.R into each other, ^ L , R  P rl>L,R{x) P1 = yoil>R,L(xP)
.
(2.23)
Thus any effect, like the weak interaction, which treats lefthanded and righthanded fermions differently will lead inevitably to parityviolating phenomena.
1—3 Symmetries and near symmetries A symmetry is said to arise in nature whenever some change in the variables of a system leaves the essential physics unchanged. In field theory the dynamical variables are the fields, and symmetries describe invariances under transformations of the fields. For example one associates with the spacetime translation ^  ^ x ^ + a^a transformation of the field x/)(x) to if){x + a). In turn, the 'essential physics' is best described by an action, at least in classical physics. If the action is invariant, the equations of motion, and hence the classical physics, will be unchanged. The invariances of quantum physics are identified by consideration of matrix elements, or equivalently, of the path integral. We begin the study of symmetries here by exploring several lagrangians which have invariances and by considering some of the consequences of these symmetries.
Noether currents The classical analysis of symmetry focusses on the lagrangian, which in general is a Lorentzscalar function of several fields, denoted by y?j, and their first derivatives d^ipi, i.e. C = C((fi, d^cpi). Noether's theorem states that for any invariance of the action under a continuous transformation of the fields, there exists a classical charge Q which is timeindependent [Q = 0) and is associated with a conserved current, d^J^ — 0. This theorem covers both internal and spacetime symmetries. For most* internal symmetries, the lagrange density is itself invariant. Given a continuous field transformation, one can always consider an infinitesimal transformation V'i(x) = tpi(x) + efi(M, A,] 6 = 1/1^,0 .
(4.8)
By direct substitution we find Fp, = dpA,,  duAp .
(4.9)
It follows from Eq. (4.7) and Eq. (4.9) that the field strength F^v is invariant under gauge transformations. A gaugeinvariant lagrangian containing a complex scalar field (p and a spin 1/2 field ip, chiral or otherwise, has the form F^F
+ ( D ^ ) D ^ + i ^ ^ + ... ,
(4.10)
where the ellipses stand for possible mass terms and nongauge field interactions. There is no contribution corresponding to a gauge boson mass. Such a term would be proportional to A^A^, which is not invariant under the gauge transformation, Eq. (4.7).
Nonabelian case The above reasoning can be generalized to nonabelian groups [YaM 54]. First we need a nonabelian group of gauge transformations and a set of fields which forms a representation of the gauge group. Then we must An abelian group is one whose elements commute. A nonabelian group is one which is not abelian.
16
/ Inputs to the Standard Model
construct an appropriate covariant derivative to act on the fields. This step involves introducing a set of gauge bosons and specifying their behavior under the gauge transformations. Finally the gauge field strength is obtained from the commutator of covariant derivatives, at which point we can write down a gaugeinvariant lagrangian. Consider fields 9 = {6*} (i — 1,... , r) which form an rdimensional representation of a nonabelian gauge group Q. The O^ can be boson or fermion fields of any spin. In the following it will be helpful to think of G as an rcomponent column vector, and operations acting on © as r x r matrices. We take group Q to have a Lie algebra of dimension n, so that the numbers of group generators, group parameters, gauge fields, and components of the gauge field strength are each n. We write the spacetimedependent group parameters as the ndimensional vector a = {aa(x)} (a = 1,..., n). A gauge transformation on O is G; = U(a)e ,
(4.11)
where the r x r matrix U is an element of group Q. For those elements of Q which are connected continuously to the identity operator, we can write U(a) = exp(ia a G a ) ,
(4.12)
where G = {G a } (a = 1,..., n) are the generators of the group Q expressed as hermitian r x r matrices. The set of generators obeys the Lie algebra [Gl^
. (1.12)
Choosing the parameter a so that Im (M'e2ta) = 0 and defining m = Re(M'e2ta), we see that £gen reduces to Cem which appears in Eq. (1.1). (iii) Renormalizability and [7(1)  Renormalizability plays a role in the preceding discussion because [7(1) symmetry by itself would admit a larger set of interaction terms. In principle, J7(l) invariant terms like ^a^F^, WF^F^, ^YYl^^FnvFap, etc. could appear in the QED lagrangian. However, they do not because the condition of renormalizability admits only those contributions which have dimension d < 4. As discussed in App. C3, the canonical dimension of boson and fermion fields is d = 1, 3/2 respectively, and each derivative adds a unit of dimension. Accordingly, the above candidate operators have d = 5, 7, 7 and ^a^F^Fa^, thus are ruled out. There remains an operator, F^F**" = which is gaugeinvariant and has dimension 4. A noteworthy aspect of this quantity is that, unlike the other operators encountered thus far, it is odd under CP. This follows from writing it as —8E • B and realizing that under CP, E —• E and B —• —B. However, a simple exercise shows that we can identify this operator as a fourdivergence F^VF^V = where K» = £e» UOL(3AvdaAp. Thus a contribution proportional to F
IIl Quantum Electrodynamics
27
can be of no physical consequence. Upon integration over spacetime, it becomes a surface term evaluated at infinity. There is nothing in the structure of QED which would cause such a surface term to be anything but zero. QED to one loop The perturbative expansion of QED is carried out about the free field limit, and is interpreted in terms of Feynman diagrams. Two distinct phenomena are involved, scattering and renormalization. The latter encompasses both an additive mass shift for the fermion (but not for the photon) and rescalings of the charge parameter and of the quantum fields. To carry out the calculational program requires a quantum lagrangian £QED t° establish the Feynman rules, a regularization procedure to interpret divergent loop integrals, and a renormalization scheme. One can develop QED using either canonical or pathintegral methods. In either case a proper treatment necessitates modification of the classical lagrangian. As we have seen, the [7(1) gauge symmetry implies a certain freedom in defining the A/i(x) field. Regardless of the quantization procedure adopted, this freedom can cause problems. For canonical quantization, the procedure of selecting a complete set of coordinates and their conjugate momenta is upset by the freedom to gauge transform away a coordinate at any given time. For path integration, the integration over gauge copies of specific field configurations gives rise to specious divergences (cf. App. A6). In either case, superfluous gauge degrees of freedom can be eliminated by introducing an auxiliary condition which constrains the gauge freedom. There are a variety of ways to accomplish this. The one adopted here is to employ the following gaugefixed lagrangian, £QED
=
]F2
  L ( 0 • A)2 + ^ {ip eofi
mo)ip ,
(1.13)
where eo and mo are respectively the fermion charge and mass parameters. The quantity £o is a realvalued, arbitrary constant appearing in the gaugefixing term. This term is Lorentzinvariant but not U(l) invariant. One of its effects is to make the photon propagator explicitly dependent on £o The value £o — 1 corresponds to Feynman gauge, whereas the limit £o —> 0 defines the Landau gauge. The zero subscripts on the mass, charge, and gaugefixing parameters denote that these bare quantities will be subject to infinite renormalizations, as will the quantum fields. This process is characterized in terms
28
/ / Interactions of the Standard Model
of quantities Zi and = Zxfyr ,
A^ = Z\I2AI ,
e 0 = ZiZ 2  1 Z 3 " 1/2 e ,
m0 = m  6m ,
(1.14)
where the superscript V labels renormalized fields. The renormalization constants Zi, Z2, and Z3 (associated respectively with the fermionphoton vertex, the fermion wavefunction, and the photon wavefunction) and the fermion mass shift 6m are chosen order by order to cancel the divergences occurring in loop integrals. For vanishing bare charge eo = 0, they reduce to Zi?2,3 = 1, 6m = 0. The Feynman rules for QED are: fermionphoton
vertex.
p
a

(1.15)
fermion propagator iSap(p): P
i (p + rao)Qjg
p
p2 — raj + ie
photon propagator

»»
a
(1.16) iD
fJ>u
(q):
(1.17) In the above e is an infinitesimal positive number. The remainder of this section is devoted to a discussion of the oneloop radiative correction experienced by the photon propagator.* Throughout, we shall work in Feynman gauge.
Fig. II—1 The full photon propagator as an iteration. * We shall leave calculation of the fermion selfenergy to Prob. II—3 and analysis of the photonfermion vertex to Sect. Vl.
IIl Quantum Electrodynamics
29
Let us define a proper or oneparticle irreducible {IPI) Feynman graph such that there is no point at which only a single internal line separates one part of the diagram from another part. The proper contributions to photon and to fermion propagators are called selfenergies. The point of finding the photon selfenergy is that the full propagator iD'^v can be constructed via iteration as in Fig. IIl. Performing a summation over selfenergies, we obtain iD' = —i
where the proper contribution
V
2 0
(1.19)
is called the vacuum polarization tensor. It is depicted in Fig. II2(a) (along with corrections to the photonfermion vertex and fermion propagator in Figs. II2(b)(c)), and is given to lowest order by
=He)2 / A T TV y ^
10!^1
1 •
(1.20) This integral is divergent due to singular high momentum behavior. To interpret it and other divergent integrals, we shall employ the method of dimensional regularization [BoG 72, 'tHV 72, Le 75]. Accordingly, we consider UaP(q) as the fourdimensional limit of a function defined in d spacetime dimensions. Various mathematical operations, such as summing over Lorentz indices or evaluating loop integrals, are carried out in d dimensions and the results are continued back to d = 4, generally expressed as an expansion in the variable* e = 4 — d. Formulae relevant to this procedure are collected in App. C4. For all theories
pq (a)
(b)
(c)
Fig. II—2. Oneloop corrections to (a) photon propagator, (b) fermionphoton vertex, and (c) fermion propagator. We shall follow standard convention is using the symbol e for both the infinitesimal employed in Feynman integrals and the variable for continuation away from the dimension of physical spacetime.
30
// Interactions of the Standard Model
described in this book, we shall define the process of dimensional regularization such that all parameters of the theory (such as e2) retain the dimensionality they have for d = 4. In order to maintain correct units while dimensionally regularizing Feynman integrals, we modify the integration measure over momentum to (1.21) The parameter ji is an arbitrary auxiliary quantity having the dimension of a mass. It appears in the intermediate parts of a calculation, but cannot ultimately influence relations between physical observables. Indeed, there exist in the literature a number of variations of the extension to d ^ 4 dimensions. These are able to yield consistent results because one is ultimately interested in only the physical limit of d = 4. Let us now return to the photon selfenergy calculation to see how the dimensional regularization is implemented. The selfenergy of Eq. (1.20), now expressed as an integral in d dimensions, is
2
e
J (2?r)d
]
)
(1.22) where we retain the same notation 11°^(g) as for d — 4 and we have already computed the trace. Upon introducing the Feynman parameterization, Dirac relations, and integral identities of App. C4, we can perform the integration over momentum to obtain
f I
x(l  x) /
9
9/1
\\*/9
'
[ml  qlx(\  x)flA (1.23) We next expand Tla^(q) in powers of e and then pass to the limit e —• 0 of physical spacetime. In doing so, we use the familiar ae = l + eln a + ... , (1.24) /o Jo
and take note of the combination r(e/2)
2
^
where 7 = 0.57221... is the Euler constant. The presence of e~l makes it necessary to expand all the other edependent factors in Eq. (1.23) and to take care in collecting quantities to a given order of e. To order e2, the
II1 Quantum Electrodynamics
31
vacuum polarization in Feynman gauge is then found to be
6TT 2
(1.26) The above expression is an example of the general property in dimensional regularization that divergences from loop integrals take the form of poles in e. These poles are absorbed by judiciously choosing the renormalization constants. Renormalization constants can also have finite parts whose specification depends on the particular renormalization scheme employed. One generally adopts a scheme which is tailored to facilitate comparison of theory with some set of physical amplitudes. In the minimal subtraction (MS) renormalization, the Z{ subtract off only the epoles, and thus have the very simple form, ^T
(* = 1.2,3).
(127)
n=l
Because the < Z\ H have no finite parts, they are sensitive only to the ultraviolet behavior of the loop integrals, and the c^n are independent of mass. The simple appearance of the MS scheme is somewhat deceptive since further (finite) renormalizations are required if the mass and coupling parameters of the theory are to be asociated with physical masses and couplings. A related renormalization scheme is the modified minimal subtraction (MS) in which renormalization constants are chosen to subtract off not only the epoles but also the omnipresent term ln(47r) — 7 of Eq. (1.25). Minimal subtraction schemes are typically used in QCD where, due to the confinement phenomenon (cf. Sect. II—2), there is no natural renormalization scale that could naturally be associated with the mass of a freely propagating quark. Yet another approach is the onshell (os) renormalization, where the renormalized mass and coupling parameters of the theory are arranged to coincide with their physical counterparts.
32
/ / Interactions of the Standard Model Onshell renormalization of the electric charge
The renormalization scale for electric charge is set by experimental determinations typically involving solid state devices like Josephson junctions. These refer to probes of the electromagnetic vertex — eT J/(p2,pi) of Fig. II2(b) with onshell electrons (p2 = Pi = m e 2 ) an
fiR
is not exactly the same quantity as the running coupling constant defined in Eq. (1.36), differing by a (small) finite renormalization. For example, the electron contribution to the running coupling in the range ml < ji 2R < M$y is
°\A)\^mi
 «" V S )l i = M s, = ^ In ^ f ,
(1.44)
which contains the dominant logarithmic dependence, but differs from Eq. (1.38) by a small additive term. However, complete calculations of * Note that the renormalization point fiR and the scale factor /x in dimensional regularization need not be identical. They are sometimes confused in the literature, and hence we use a different notation for the two quantities.
36
/ / Interactions of the Standard Model
all corrections to physical observables using the two schemes will yield the same answer. Since the running coupling constant is but a bookkeeping device, one's choice is a matter of taste or of convenience. Regardless of the specific definition employed for a(# 2 ), we see that as the energy scale is increased (or as distance is decreased), the running electric charge grows, as is expected on physical grounds from vacuum polarization. The use of a massindependent scheme is convenient for identifying the high energy scaling behavior of gauge theories. One useful feature is in the calculation of the oneloop beta function. Dimensional analysis requires that the oneloop charge renormalization be of the form, 9 = 90
 9ob ( Z~2 )
( +
finite t e r m s
(1.45)
)
where g is the 'charge' associated with the gauge theory being considered. Choosing the renormalization point as q2 = —^ 2R and forming the beta function as in Eq. (1.42), we see that (3 = bg3. This allows the beta function to be simply identified with the coefficient of e" 1 to this order. II—2 Quantum Chromodynamics Chromodynamics, the nonabelian gauge description of the strong interactions, contains quarks and gluons instead of electrons and photons as its basic degrees of freedom [PrG 72, MaP 78]. A hallmark of quantum chromodynamics is asymptotic freedom [GrW 73a,b, Po 73], which reveals that only in the shortdistance limit can perturbative methods be legitimately employed. The necessity to employ approaches alternative to perturbation theory for longdistance processes motivates much of the analysis in this book. SU(3) gauge symmetry Chromodynamics is the SU(3) nonabelian gauge theory of color charge. The fermions which carry color charge are the quarks, each with field t 3L m , where a = u, d, s,... is the flavor label and j = 1,2,3 is the color index. The gauge bosons, which also carry color, are the gluons, each with field Aa, a = 1,..., 8.* Classical chromodynamics is defined by the lagrangian
£color = ~\ F?F%, + J2 ^VttVjk ~ ™
(Q)
Wia) ,
(21)
* In this section, it will be particularly important to explicitly display color indices. We shall reserve indices which begin the alphabet for gluon color indices (e.g., a,b,c — 1,...,8 ), use midalphabetic letters for quark color indices (e.g., j,k,£ = 1,2,3 ), and employ greek symbols for flavor indices.
II2 Quantum Chromodynamics
37
where the repeated color indices are summed over. strength tensor is F% = d^Al  dvAl  9ifahcA\Al
The gauge field
,
(2.2a)
gs is the SU(3) gauge coupling parameter, and the quark covariant derivative is T>^ = (dfl + igsAa^W
.
(2.2b)
The lagrangian of Eq. (2.1) is invariant under local SU(3) transformations of the color degree of freedom, under which the quark and gluon fields transform as given earlier in Eqs. (14.11), (14.17). Equations of motion for the quark and gluon fields are
(2 3)
*
In its quantum version, the gs —• 0 limit of £coior describes an exceedingly simple world. There exist only free massless spin one gluons and massive spin onehalf quarks. However, the full theory is quite formidable. In particular, accelerator experiments reveal a particle spectrum which bears no resemblance to that of the noninteracting theory. The group SU{3) has an infinite number of irreducible representations R. The first several are R = 1, 3, 3*, 6, 6* 8, 10, 10*, . . . , where we label an irreducible representation in terms of its dimensionality. Quarks, antiquarks, and gluons are assigned to the representations 3, 3*, 8 respectively. We denote the group generators for representation R by {Fa(R)} (a = 1 , . . . , 8). The quantities A/2 are group generators for the d = 3 fundamental representation, i.e., F(3) = A/2. They have the matrix representation 0 0 A4 = . 0 0 1 0 A2 = \ i
0
0
A5 =
0 0 0 0 i 0 0
A3 = I 0
1
0 I
A6 =  0 u 0
0 0u
i> 0 .
^
=
/ —i 0 0 1I j
73 0 0
.
1 0 /
0
73 0
0 0 2
73(2 4^ y
'
J
As generators, they obey the commutation relations [Aa, A6] = 2i/ a6c A c
(a, b, c = 1 , . . . , 8)
(2.5a)
38
/ / Interactions of the Standard Model
where the /coefficients are totally antisymmetric structure constants of SU(3). There exist corresponding anticommutation relations {Aa, Xb} = pab I + 2dabc\c
(a, 6, c = 1,..., 8)
(2.5b)
with dcoefficients which are totally symmetric. Values for fabc and dabc are given in Table II—1. Useful trace relations obeyed by the {Aa} are Tr Aa = 0
(a = l , . . . , 8)
(2.6)
from Eq. (2.4) and TV \aXb = 26ab (a, b = 1,..., 8) from Eq. (2.5). The statement of completeness takes the form, KjKi = —SijSid + 26aSjk
(t, j ,fc,J = 1,2,3) ,
(2.7) (2.8)
where a = 1,..., 8 is summed over. Useful labels for the irreducible representations of SU(3) are provided by the Casimir invariants. For any representation R, the quadratic Casimir invariant C2(R) is defined by squaring and summing the group generators {Fa(i?)}5 8
o=l
There is also a thirdorder Casimir invariant, 8
CS(R)I=
] T dabcFa(R)Fb(R)Fc(R) a,6,c=l
Table II—1. Nonvanishing /, d coefficients abc
123 1 147 1/2 156 1/2 246 1/2 257 1/2 345 1/2 367 1/2 458 V3/2 678 V3/2
dabc
abc
dabc
118 l/>/3 355 1/2 146 1/2 366 1/2 157 1/2 377 1/2 228 1/VS 448 l/2>/3 247 1/2 558 l/2>/3 256 1/2 668 l/2\/3 338 1/V3 778 l/2\/3 344 1/2 888 1/V5
.
(2.10)
II2 Quantum Chromodynamics
39
The quark and antiquark states form the bases for the smallest nontrivial irreducible representations of SU(3). It is possible to use products of them, say p factors of quarks and q factors of antiquarks, to construct all other irreducible tensors in SU(S). Each irreducible representation R is then characterized by the pair (p, q). For example, we have the correspondences 1 ~ (0,0), 3 ~ (1,0), 3* ~ (0,1), 8 ~ (1,1), 10 ~ (3,0), etc. The (p, q) labeling scheme provides useful expressions for the dimension of a representation, dip, q) = (p + l)(q + l)(p + q + 2)/2 ,
(2.11)
and of the two Casimir invariants, C2(p,