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provided by CERN Document Server hep-ph/9512224 TUM–HEP–227/95 November 1995
The Goldstone boson equivalence theorem with fermions
Loyal Durand∗
Department of Physics, University of Wisconsin-Madison, 1150 University Avenue, Madison, WI 53706
and
Kurt Riesselmann†
Physik Department T30, Technische Universit¨at M¨unchen, James-Franck Strasse, 85747 Garching b. M¨unchen, Germany
Abstract The calculation of the leading electroweak corrections to physical transition matrix 2 2 2 2 2 elements in powers of MH /v can be greatly simplified in the limit MH MW ,MZ through the use of the Goldstone boson equivalence theorem. This theorem allows the vector bosons W ± and Z to be replaced by the associated scalar Goldstone bosons w±, z which appear in the symmetry breaking sector of the Standard Model in the limit of vanishing gauge couplings. In the present paper, we extend the equivalence theorem systematically to include the Yukawa interactions between the fermions and the Higgs and Goldstone bosons of the Standard Model. The corresponding Lagrangian LEQT is given, and is formally renormalized to all orders. The renormalization conditions are formulated both to make connection with physical observables and to satisfy the requirements underlying the equivalence theorem. As an application of this framework, we calculate the dominant radiative corrections to fermionic Higgs decays at one loop including the virtual effects of a heavy top quark. We apply the result to the decays H → tt¯and H → b¯b, and find that the equivalence theorem results including fermions are quite accurate numerically for Higgs-boson masses MH > 400 (350) GeV, respectively, even for mt = 175 GeV.
∗Electronic address: [email protected] †Electronic address: [email protected]
1 1 Introduction
The physics of the Higgs-sector is one of the most important subjects of high-energy experi- ments in the near future. To distinguish the Standard Model (SM) Higgs-sector from other theoretical models, the calculation of radiative corrections to high-energy processes may be very important. In addition, the effects of a Higgs boson on low-energy observables such as the ρ-parameter already need to be included in present precision measurements. However, the calculation of the complete higher-order corrections is a difficult task. It is therefore impor- tant if one can calculate the most significant corrections relatively simply within a well-defined approximation. In the case of a heavy Higgs boson, such an approximation is provided by the equivalence theorem (EQT) [1, 2]. This theorem shows that the leading correction to a 2 2 physical process in powers of MH /v can be obtained at any order by replacing the vector bosons W ± and Z of the SM by the associated scalar Goldstone bosons w±, z which appear in the symmetry breaking sector of the theory in the limit of vanishing gauge couplings. This results in a substantial simplification of the calculations. The equivalence theorem is known to hold for MH MW ,MZ. An important question, however, is how heavy the Higgs boson must be before the equivalence theorem gives a good approximation to the full electroweak theory. The existence of a heavy top quark leads to a large Yukawa coupling gt in the SM. In the case of a not-so-heavy Higgs boson, one can expect the radiative corrections which involve gt to be comparable in magnitude to the corrections which involve the quartic Higgs coupling λ = 2 2 MH /2v . It is therefore desirable to take these contributions into account, and to calculate them in a simple way. This can again be done using the equivalence theorem: in the limit of vanishing gauge couplings, the fermions couple only to the Higgs boson and the Goldstone bosons in the symmetry breaking sector of the theory. However, the specific calculational scheme used has to be consistent both with the EQT as formulated for the bosons [1, 2], and with the complete SM. In the present paper we show how this can be achieved by extending the equivalence theorem, including fermions, and implementing consistent renormalization conditions. The resulting framework is a foundation for future calculations of leading higher- order electroweak corrections, and it is applicable to all orders in λ and gf . Gauge interactions are explicitly set to zero. Previous calculations based on the EQT have mostly been concerned with processes in which both the external and internal particles are Higgs bosons H or longitudinally polarized gauge bosons WL and ZL [1,2,3,4,5,6,7,8,9,10],andhaveneglectedcorrectionsfrom 4 internal fermion loops. An exception is the work of Barbieri et al. [11], who calculated the mt corrections to the decays Z → µ+µ− and Z → b¯b using a method based on the equivalence theorem, but with the fermion fields renormalized only to one loop. The framework presented here allows the systematic inclusion of internal fermion loops at higher orders, as well as the calculation of radiative corrections to processes involving external fermions. We illustrate our method at the one-loop level using the fermionic decay of the Higgs boson, H → ff¯. The equivalence theorem result approximates the full electroweak correction to Γ(H → tt¯) very well for MH > 400 GeV, but fails near the threshold at 350 GeV for mt = 175 GeV, where the effects of the gauge interactions become important. The result for Γ(H → b¯b) is also quite good for MH > 350 GeV, though the accuracy is reduced by a cancellation between the Yukawa and Higgs-boson contributions. The gauge couplings give the dominant, very small, corrections for lower values of MH.
2 2 The Lagrangian and renormalization scheme 2.1 Framework We will be concerned in later sections with the calculation of the leading contributions to the decay rate of the Higgs boson to fermions, H → ff¯, in the limit of large Higgs-boson mass, MH MW . Our tool for extracting the leading contributions in powers of MH /MW will be the Goldstone boson equivalence theorem [1, 2, 12, 13, 14, 15]. This theorem states that the dominant electroweak contribution to a Feynman graph can be calculated by replacing the gauge bosons W ±, Z of the full electroweak theory by the would-be Goldstone bosons w±, z of the symmetry-breaking sector of the theory, and ignoring the gauge couplings. The result is an 2 expansion for the dominant contributions in powers of GF MH , or equivalently, in powers of the quartic Higgs coupling λ. In the limit g, g0 → 0, the tranverse gauge bosons decouple from the fermions and the scalar fields, and can be ignored. The remainder of the standard electroweak model reduces to a theory defined by the equivalence-theorem Lagrangian LEQT = LH +LF , where LH is the Lagrangian for the scalar fields in the symmetry-breaking sector, and LF is the Lagrangian for the fermions which includes their Yukawa interactions with the scalar fields. This reduced theory gives a good approximation to the full electroweak theory for 2 2 MW MH , provided that the couplings are properly defined and the Goldstone modes are 2 2 renormalized at a momentum scale p MH [13]. We will work entirely in the equivalence theorem limit using LEQT ,
0 0 LEQT = LH + LF +counterterms, (1) where
0 1 † µ 1 † 2 1 2 † LH = 2 (∂µΦ) (∂ Φ) − 4 λ Φ Φ + 2 µ Φ Φ , (2) 0 ¯ ¯ ¯ LF = iΨL 6∂ ΨL + iψt,R 6∂ψt,R + iψb,R 6∂ψb,R √1 g Ψ¯ Φ˜ ψ √1 g Ψ¯ Φ ψ +h.c.+ . (3) − 2 t L t,R − 2 b L b,R ···
0 We have written only the top- and bottom quark contributions to LF ; the remaining fermionic contributions have the same form, with no righthanded contributions for the neutrinos. In these expressions, ΨL,Φ,andΦareSU(2)˜ L doublets with normalizations defined by
ψt,L 1 5 ΨL = ,ψf,L = 2(1 − γ ) ψf , (4) ψb,L √ w+iw 2w+ Φ= 1 2 = , (5) h + iz h + iz h iz ˜ ∗ h − iz √− Φ=iσ2Φ = = − . (6) −w1 + iw2 − 2w
The righthanded fields are SU(2)L singlets, with
1 5 ψf,R = 2(1 + γ ) ψf . (7)
The expressions in Eqs. (2) and (3) include all possible SU(2)L×U(1)Y symmetric terms consistent with renormalizability of the Lagrangian. The counterterms necessary to effect
3 the renormalization without breaking the symmetry must have the same forms, and can be introduced by independent multiplicative rescalings of each term above. Thus, in the case of the Higgs Lagrangian, all possible symmetric counterterms are generated by multiplying the 2 2 2 kinetic energy term by a factor ZΦ, replacing λ by λ + δλ,andµ by µ + δµ . Because the minimum in the Higgs potential is at a nonzero value of Φ†Φforµ2 >0, Φ has a nonzero expectation value in the physical vacuum,
h Ω | Φ†Φ | Ω i = v2 . (8)
We will absorb the vacuum expectation value by rewriting the field h as h → H + v.This results in an expansion around the physical vacuum in which
h Ω | H | Ω i = h Ω | w | Ω i =0. (9)
The renormalized Higgs Lagrangian then has the form
1 µ µ 1 2 2 2 LH = 2 ZΦ (∂µw · ∂ w + ∂µH∂ H)−4 (λ+δλ) w + H +2vH 1 2 2 2 2 2 − 2 (λ + δλ) v − µ − δµ w +H +2vH , (10) where w is the SO(3) vector (w1,w2,w3)with w3 =z, and we have dropped an additive constant. It is convenient for calculation to define a set of rescaled or “bare” fields H0, w0 and the corresponding SU(2)L doublet Φ0 such that the kinetic terms in LH have the customary unit normalization, 1/2 1/2 1/2 H0 = ZΦ H, w0 =ZΦ w, Φ0 =ZΦ Φ, (11) and to introduce a corresponding bare vacuum expectation value v0, a bare coupling λ0,and 2 a parameter δm0,
1/2 v0 = ZΦ v, (12) λ δλ λ = 1+ , (13) 0 Z2 λ Φ 2 −1 2 2 2 δm0 = ZΦ µ + δµ − (λ + δλ) v . (14)
With this rescaling, we obtain the form of LH that we will use,
1 µ µ 1 2 2 2 LH = 2 (∂µw0 · ∂ w0 + ∂µH0 ∂ H0) − 4 λ0 w0 + H0 +2v0H0 1 2 2 2 +2δm0 w0 + H0 +2v0H0 . (15)
The fermion Lagrangian L F can be treated in a similar fashion. There is a separate SU(2)L×U(1)Y -symmetric counterterm for each term in Eq. (3). These counterterms can be absorbed in the definitions of a set of bare fields and couplings to bring LF into the form
¯ 0 0 ¯0 0 ¯0 0 LF = iΨL 6∂ ΨL + iψt,R 6∂ψt,R + iψb,R 6∂ψb,R √1 g0 Ψ¯ 0 Φ˜ ψ0 √1 g0 Ψ¯ 0 Φ ψ0 +h.c., (16) − 2 t L 0 t,R − 2 b L 0 b,R
4 where
0 1/2 ΨL = ZL ΨL , (17) 0 1/2 ψf,R = Zf,R ψf,R ,f=t, b, (18) g δg g0 = f 1+ f . (19) f 1/2 gf ZΦ
Note that there are separate renormalization constants for the ψt,R and ψb,R,andasingle −1/2 0 renormalization constant for the left-handed doublet ΨL [16]. The factor ZΦ in gf has been introduced for later convenience.
2.2 Fixing the renormalization of the Higgs Lagrangian
In order that the theory defined by LEQT correspond to the equivalence theorem limit of the standard electroweak model, the definitions of the couplings and the renormalization scheme must be chosen to be consistent with that limit. Such a choice is not automatic. It is necessary ± 2 2 2 that: (i), the w and z fields be renormalized at a momentum scale p with |p |MH [13]; and (ii), the couplings be defined in terms of physical standard-model quantities which have well-defined limits for g, g0 → 0. In the following, we will use an on-mass-shell renormalization scheme, define the quartic Higgs coupling in terms of the Fermi constant GF and the physical Higgs-boson mass MH , and relate the Yukawa couplings to the physical masses of the fermions. The w± and z bosons are guaranteed to be massless by the Goldstone theorem [17]. We ± 2 will therefore renormalize the w0 and z0 fields at p = 0, thus satisfying condition (i), and 2 2 will renormalize the Higgs field at p = MH . The quartic coupling λ will be defined [18] to be given exactly by the relation √ 2 2 2 λ = MH /2v = GF MH / 2, (20) where GF is the Fermi constant obtained from the muon decay rate using the standard electromagnetic radiative corrections, and v is the physical vacuum expectation value, v = −1/4 −1/2 2 GF . While this definition involves a process at a low momentum transfer, it connects smoothly with equivalence theorem limit as shown in [18, 19]. The determination of the renormalization constants proceeds as follows. The real part of the two-point function or inverse propagator Γ(2)(p2) for each of the particles must vanish for p2 equal to the square of the physical mass of that particle, Γ(2)(m2) = 0. The two-point functions calculated using the bare fields are easily seen to have the form
(2) 2 2 0 2 2 Γw0 (p )=p−Πw(p)+δm0 , (21) (2) 2 2 0 2 2 Γz0 (p )=p−Πz(p)+δm0 , (22) Γ(2)(p2)=p2 Π0(p2) 2λv2+δm2 , (23) H0 −H − 00 0 where the Π’s are the bare self-energy functions. Since there is only a single mass counterterm 2 δm0 in the Higgs Lagrangian, Eq. (15), the vanishing of the renormalized masses mw and mz required by the Goldstone theorem leads to the relations
2 0 0 δm0 =Πw(0) = Πz(0) , (24)
5 2 ± and determines δm0. The self-energy functions for the bare w and z fields must therefore be equal at p2 = 0, an identity that holds to all orders in perturbation theory. This would not be surprising in the absence of Yukawa couplings since the Goldstone fields then have an SO(3) 0 2 0 2 2 symmetry, and Πw(p )andΠz(p) are necessarily identical for all values of p . However, the presence of fermions with unequal Yukawa couplings in LEQT breaks the SO(3) symmetry, and destroys the equality of the self-energy functions except at the special point p2 =0[20]. 0 0 The equality Πw(0) = Πz(0) can be checked to one loop using the results we give later. In 2 0 the following, we will replace δm0 by Πw(0). 2 2 The bare coupling λ0 can be determined by using the definition λ = MH /2v and Eq. (23). The on-mass-shell condition for the Higgs boson, Γ(2)(M 2 ) = 0, gives the relation H0 H
2 0 2 2 MH =2λ0v0 +ΠH(MH)−δm0. (25) 2 2 2 2 0 Upon replacing v0 by ZΦv = ZΦMH /2λ and δm0 by Πw(0) in this expression and solving for λ0, we find that [21] δλ 1 λ ReΠ0 (M 2 ) − Π0 (0) λ = λ 1+ = 1 − H H w . (26) 0 λ Z2 Z M 2 Φ Φ H The condition that v be the physical vacuum expectation value requires the vanishing of the truncated one-point function for the Higgs field,
Γ(1)(0) = iT + iv δm2 =0, (27) H0 − 0 0 0 where T0 is the sum of all Higgs tadpole graphs (see Fig. 1) calculated using the bare fields. 2 2 This gives the further relation δm0 = T0/v0,soδm0 can be calculated either as a self energy or as a tadpole contribution. The resulting identity,
0 Πw(0) = T0/v0, (28) provides a useful check on the calculations [22]. The vanishing of the one-point function, Eq. (27), implies that the tadpole diagrams and the tadpole counterterm in Eq. (15) cancel order-by-order in the perturbation expansion and can be dropped together, as discussed by Taylor [23]. We will adopt this simplification in the following calculations. Finally, the wave function renormalization constants Zw , Zz,andZH which relate the ± bare fields w0 , z0, H0 to physical fields,
± 1/2 ± 1/2 1/2 w0 = Zw wphys,z0=Zzzphys,H0=ZHHphys, (29) are determined by the condition that the propagators for the physical fields have unit residue at the particle poles. The bare propagators do not, but instead have residues Zi given by
−1 d (2) 2 Zi = 2 Γi (p ) . (30) dp p2=m2 i !
Thus, using Eqs. (21-23), 1 d =1− Π0 (p2) , (31) Z dp2 w w p2=0
6 1 d 0 2 =1− 2 Πz (p ) , (32) Z dp 2 z p =0
1 d 0 2 =1− 2 ΠH (p ) . (33) ZH dp p2=M 2 H
In the presence of fermions with different Yukawa couplings or masses, e.g., top and bottom 0 0 quarks with mt 6= mb, the self-energy functions Πw and Πz are not equal except at the isolated 2 point p = 0, and, as a result, Zz 6= Zw. The single wave function renormalization constant ZΦ introduced earlier is sufficient, along with the other renormalization constants, to absorb the divergences in the Higgs sector of the theory. The constants Zw , Zz ,andZH can therefore differ from ZΦ only by finite multiplica- tive factors Z˜i, Zi = Z˜iZΦ . (34)
The problem which remains is that of determining ZΦ,anessentialstepinconnectingthe scalar theory to the full electroweak theory through the equivalence theorem. The crucial observation is that the Fermi constant GF , and therefore the vacuum expectation value v,is defined through charged-current processes, i.e., muon decay and superallowed nuclear beta decays. These involve the W rather than the Z boson. The connection can then be made using the Ward identities for the electroweak charged current which underlie the equivalence theorem [13, 11]. In particular, Bagger and Schmidt [13] show that W ± scattering amplitudes calculated in the full electroweak theory are related to those calculated using the equivalence theorem by powers of the ratio [24]
0 1/2 1/2 1/2 MW ZW g0v0 ZW ZΦ 2 C = 1/2 = 1/2 = 1/2 1+O(g ) . (35) MW Zw gv Zw Zw They then use the Ward identity to establish that C = 1 up to corrections of order g2 under conditions satisfied by our renormalization scheme. Thus, ZΦ = Zw and Z˜w = 1 in the limit g, g0 → 0.
2.3 Fixing the renormalization of the fermion Lagrangian
When written with the SU(2)L spinor products expanded, the fermion Lagrangian LF in Eq. (16) assumes the form
¯0 0 ¯0 0 ¯0 0 ¯0 0 LF = iψt 6∂ψt −(mt +δmt)ψt ψt − iψb 6∂ψb −(mb +δmb)ψb ψb 0 0 0 0 0 5 0 0 0 0 0 0 5 0 √1 g ψ¯ H ψ + √i g ψ¯ γ z ψ √1 g ψ¯ H ψ √i g ψ¯ γ z ψ − 2 t t 0 t 2 t t 0 t − 2 b b 0 b − 2 b b 0 b 0 ¯0 + 0 0 ¯0 − 0 0 ¯0 + 0 0 ¯0 − 0 +gt ψt,R w0 ψb,L + gt ψb,L w0 ψt,R − gb ψb,R w0 ψt,L − gb ψt,L w0 ψb,R , (36) where we have used the relations in Eqs. (11), (12), and (19). The parameters mf and gf are defined to be the physical masses and Yukawa couplings of the quarks while δmf is the mass or coupling counterterm, v δgf mf = gf √ ,δmf=mf . (37) 2 gf Our calculations will be carried out using the bare fields and the standard free propagators 0 for the quarks, and yield corrected fermion propagators iSF of the form
0 i iSF = 0 . (38) 6p − mf − Σf − δm
0 The self-energy Σf contains an axial vector component as well as the vector and scalar com- ponents familiar in parity-conserving theories. In particular, suppressing the quark label,
0 0 2 5 0 2 0 2 Σ =6p ΠV (p )+ 6pγ ΠA(p )+mΠS(p ). (39)
The physical propagators iSF are related to the bare propagators through the connection between the physical and bare fields,
−1/2 −1/2 0 ψ = ψR + ψL = ZR PR + ZL PL ψ , (40) 1 5 1 5 where PR = 2 (1 + γ )andPL = 2(1 − γ ). Recalling the definition of the propagator, iSF (x − y)=hΩ|T ψ(x),ψ¯(y) |Ωi, we find that