Low-Energy Photon-Photon Collisions to Two-Loop Order

59 Pages·0.37 MB·English

by  S. Bellucci

BUTP-93/18
LNF-93/077 (P)
PSI-PR-93-17
Low-energy photon-photon collisions
4 to two-loop order ♯
9
9
1

n
a S. Belluccia, J. Gasserb and M.E. Sainiob,c,d
J

4

1
v February 2008
6
0
2
1
Abstract
0
4
9 We evaluate the amplitude for γγ π0π0 to two loops in chiral perturbation
/ →
h
theory. The three new counterterms which enter at this order in the low-
p
- energy expansion are estimated with resonance saturation. We find that the
p
e cross section agrees rather well with the available data and with dispersion
h
: theoretic calculations even substantially above threshold. Numerical results
v
i for the Compton cross section and for the neutral pion polarizabilities are also
X
given to two-loop accuracy.
r
a
♯ Work supported in part by Schweizerischer Nationalfonds and by the EEC
Human Capital and Mobility Program.
a) INFN-Laboratori Nazionali di Frascati, P.O.Box 13, I-00044 Frascati, Italy.
b) Institute for Theoretical Physics, University of Bern, Sidlerstrasse 5, CH-
3012 Bern, Switzerland.
c) Paul Scherrer Institut, CH-5232 Villigen PSI, Switzerland.
d) During the academic year 1992-93 on leave of absence from Dept. of Theo-
retical Physics, University of Helsinki, Finland.
e-mail: [e-mail]; [e-mail]; Sainio@finuhcb.bitnet
1 Introduction
The cross section for γγ π0π0 and for γγ π+π has been calculated some
−
→ →
time ago [1, 2] in the framework of chiral perturbation theory (CHPT) [3]-[7] and
of dispersion relations. For charged pion-pair production, the chiral calculation [1]
at next-to-leading order is in good agreement with the Mark II data [8] in the low-
energy region. On the other hand, for γγ π0π0, the one-loop prediction [1, 2]
→
disagrees with the Crystal Ball data [9] and with dispersion theoretic calculations
[10]-[16] even near threshold.
In the process γγ π+π , the leading contribution1 is generated by tree dia-
−
→
grams. One has a control on higher order corrections in this case, in the sense that
it is explicitly seen that the one-loop graphs do not modify the tree amplitude very
strongly near threshold [1]. Tree diagrams are absent for γγ π0π0 which starts out
→
with one-loop graphs. It is the aim of this article to establish the region of validity
of the chiral representation of this process by evaluating the amplitude at two-loop
order.
Is a next-to-leading order calculation sufficient in this case? If the corrections are
large, the reliability of the result is certainly doubtful. However, a glance at the data
shows that the corrections needed to bring CHPT and experiment into agreement
are not large–a 25-30% change in amplitude is sufficient. Corrections of this size are
rather normal in reactions where pions in an isospin zero S-wave state are present
[19]. As an example we mention the isospin zero S-wave ππ scattering length, whose
tree-level value [20] receives a 25% correction from one-loop graphs [4]. Corrections
of a similar size are present in the scalar form factor of the pion [21].
The amplitude for γγ π0π0 also describes Compton scattering on neutral pions
→
by analyticity and crossing. Do sizeable corrections in γγ π0π0 then also show
→
up in γπ0 γπ0 ? Since there are no strongly interacting particles in the final
→
state in this case, one might be led to suspect that the one-loop amplitude is a good
approximation for this reaction. We find it interesting that this is not the case–the
corrections to the leading-order term are in fact very large in this channel.
The electromagnetic polarizabilities characterize aspects of the inner structure
of hadrons. With the two-loop expression for the amplitude at hand, it is straight-
1 In this article, we denote the first nonvanishing contribution to any quantity by ”the leading-
order term”, independently of whether it starts out at tree level or at higher order in the chiral
expansion.
2
forward to evaluate the polarizabilities at next-to-leading order in the quark mass
expansion. Renormalization group arguments show that this expansion contains log-
arithmicsingularitiesofthetypeM ln2M2 andM lnM ,andanorderofmagnitude
π π π π
estimate reveals that these contributions may easily be as large as the leading-order
term, unless the relevant Clebsch-Gordan coefficient is small. We find that the latter
is the case.
Recently, a reformulation of CHPT has been given [22], where the effective la-
grangian includes into each order additional terms which in the standard CHPT
(considered here) are relegated to higher orders. To all orders, the two perturbative
schemes are identical–in each finite order, they may, however, substantially differ.
For an analysis of the process γγ π0π0 in this generalized framework we refer the
→
reader to Ref. [23].
The article is organized as follows. In section 2, we set up the notation. In section
3 we describe the low-energy expansion in a general manner and outline the specific
procedure for the two-loop case in sections 4 and 5. The low-energy constants which
occur in the amplitude for γγ π0π0 at two-loop order are determined in section
→
6. Section 7 contains a discussion of the amplitude and of the cross section at two-
loop order. The Compton amplitude and the pion polarizabilities are described in
section 8, whereas section 9 is devoted to a comparison of the chiral expansion with
the dispersive analysis of γγ π0π0 by Donoghue and Holstein [13]. Finally, a
→
summary and concluding remarks are presented in section 10.
2 Kinematics
The matrix element for pion production
γ(q )γ(q ) π0(p )π0(p ) (2.1)
1 2 1 2
→
is given by
< π0(p )π0(p )out γ(q )γ(q )in >= i(2π)4δ4(P P )TN , (2.2)
1 2 1 2 f i
| −
with
TN = e2ǫµǫνV ,
1 2 µν
V = i dxe i(q1x+q2y) < π0(p )π0(p )out Tj (x)j (y) 0 > . (2.3)
µν − 1 2 µ ν
| |
Z
3
Here j is the electromagnetic current, and α = e2/4π 1/137. The decomposition
µ
≃
of the correlator V into Lorentz invariant amplitudes reads with q2 = q2 = 0 (see
µν 1 2
appendix A)
V = A(s,t,u)T +B(s,t,u)T +C(s,t,u)T +D(s,t,u)T ,
µν 1µν 2µν 3µν 4µν
s
T = g q q ,
1µν µν 1ν 2µ
2 −
T = 2s∆ ∆ ν2g 2ν(q ∆ q ∆ ) ,
2µν µ ν µν 1ν µ 2µ ν
− − −
T = q q ,
3µν 1µ 2ν
T = s(q ∆ q ∆ ) ν(q q +q q ) ,
4µν 1µ ν 2ν µ 1µ 1ν 2µ 2ν
− −
∆ = (p p ) , (2.4)
µ 1 2 µ
−
where
s = (q +q )2, t = (p q )2, u = (p q )2 ,
1 2 1 1 2 1
− −
ν = t u , (2.5)
−
are the standard Mandelstam variables. The tensor V satisfies the Ward identities
µν
qµV = qνV = 0 . (2.6)
1 µν 2 µν
The amplitudes A and B are analytic functions of the variables s,t and u, symmetric
under crossing (t,u) (u,t). The quantities C and D do not contribute to the
→
process considered here (gauge invariance).
It is useful to introduce in addition the helicity amplitudes
H = A+2(4M2 s)B ,
++ π −
8(M4 tu)
H = π − B . (2.7)
+
− s
The helicity components H and H correspond to photon helicity differences
++ +
−
λ = 0,2, respectively. They have partial wave expansions involving even J λ,
≥
H = hJ(s)dJ (cosθ) ,
++ + 00
J=0,2,4...
X
H = hJ(s)dJ (cosθ) , (2.8)
+ 20
− −
J=2,4,6...
X
where θ is the scattering angle in the center-of-mass system, ~q p~ = ~q p~ cosθ.
1 1 1 1
· | || |
4
With our normalization of states < p~ ~p >= 2(2π)3p0δ3(p~ p~ ), the differential
1 | 2 1 1− 2
cross section for unpolarized photons in the center-of-mass system is
dσγγ π0π0 α2s
→
= β(s)H(s,t) ,
dΩ 64
H(s,t) = H 2 + H 2 ,
++ +
| | | − |
β(s) = (1 4M2/s)1/2. (2.9)
− π
The amplitude for Compton scattering
γ(q )π0(p ) γ(q )π0(p )
1 1 2 2
→
may be obtained by crossing. In the center-of-mass system, the cross section for
unpolarized photons is
dσγπ0 γπ0 α2
→ = t¯2H(t¯,s¯) , (2.10)
dΩ 16s¯
with
s¯= (q +p )2 , t¯= (q q )2 .
1 1 2 1
−
Finally, the optical theorem in the Compton channel reads with our phase convention
1
e2ImB = σγπ0(s¯) . (2.11)
|s=0,t=s¯ 4(s¯ M2) tot
− π
This relation fixes the phase of A through Eq. (2.4).
The physical region for the reactions γγ π0π0 and γπ0 γπ0 is displayed
→ →
in Fig. 1, where we also indicate with shaded lines the nearest singularities in the
amplitudes A and B. These singularities are generated by two-pion intermediate
states in the s,t and u channel.
3 Low-energy expansion
We consider QCD with two flavours in the isospin symmetry limit m = m = mˆ
u d
and equip the underlying lagrangian with hermitean, colour neutral external fields
v,a,s and p in the standard manner,
= 0 +q¯γµ(v +a γ )q q¯(s iγ p)q . (3.1)
L LQCD µ µ 5 − − 5
Here 0 denotestheQCDlagrangianatzero quarkmass, whereasmˆ iscontainedin
LQCD
the scalar field s(x). The lagrangian(3.1) is invariant under local SU(2) SU(2)
L R
× ×
U(1) transformations
1
q [(1+γ )g +(1 γ )g ]q (3.2)
5 R 5 L
→ 2 −
5
with
g = eiφV ,
R,L R,L
V SU(2) ,
R,L
∈
φ = diag(φ ,φ ) ,φ R , (3.3)
0 0 0
∈
provided that the external fields are subject to the gauge transformation
rµ′ = gRrµgR† +igR∂µgR† ,
lµ′ = gLlµgL† +igL∂µgL† ,
s′ +ip′ = gR(s+ip)gL† ,
r = v +a , l = v a . (3.4)
µ µ µ µ µ µ
−
Since the charge is not a generator of SU(2), we consider in the following the case
a = 0 , v = 0 , (3.5)
µ µ
h i h i 6
where A denotes the trace of the matrix A. The condition (3.5) is consistent with
h i
the transformation law (3.4). The Green functions of the theory are generated by
the vacuum-to-vacuum amplitude
eiZ(v,a,s,p) =< 0 0 > . (3.6)
out | in v,a,s,p
The generating functional Z admits an expansion in powers of the external momenta
and of the quark masses [3]-[5],
Z = Z +Z +Z +... , (3.7)
2 4 6
where Z denotes a term of order En. We write the corresponding expansion of the
n
amplitudes as
I = I +I +I +... ; I = Vµν,A,B , (3.8)
2 4 6
where it is understood that Z generates I 2. To calculate Vµν, we set
n n
s = mˆ1 , v = Qv¯ , p = τ3p¯ , a = 0 , (3.9)
µ µ µ
where
1
Q = diag(2, 1) (3.10)
3 −
is the charge matrix, and where v¯ and p¯denote flavour neutral external fields. Vµν
µ
is obtained from the term of order v¯2p¯2 in Z.
2Notice that I is not of order En.
n
6
3.1 Terms at order E2
In the meson sector, Z is given by the classical action
2
Z = dx (U,v,a,s,p) , (3.11)
2 2
L
Z
where is the nonlinear σ-model lagrangian
2
L
F2
= D UDµU +χ U +χU , (3.12)
2 µ † † †
L 4 h i
evaluated at the solution to the classical equation of motion δ = 0. The 2 2
2
L ×
unitary matrix U contains the pion fields, R
φ φ2
U = σ +i , σ2 + = 1 ,
F F2
π0 √2π+
φ = = φiτi . (3.13)
 √2π π0 
−
−
 
It transforms as
G
U → gRUgL† (3.14)
under G = SU(2) SU(2) U(1). The covariant derivative is
L R
× ×
D U = ∂ U ir U +iUl , (3.15)
µ µ µ µ
−
and the field χ denotes the combination
χ = 2B(s+ip). (3.16)
F is the pion decay constant in the chiral limit, F = F(1+O(mˆ)),F 93 MeV,
π π
≃
and B is related to the order parameter < 0 q¯q 0 >. The physical pion mass is
| |
M2 = M2(1+O(mˆ)) ,
π
M2 = 2mˆB. (3.17)
is referred to as the effective lagrangian at order E2.
2
L
The term of order O(v¯2p¯2) in the classical action Z vanishes and, therefore, one
2
has
A = B = Vµν = 0 . (3.18)
2 2 2
7
3.2 Higher orders in the energy expansion
At higher orders in the energy expansion, the effective lagrangian consists of a string
of terms. Reintroducing momentarily h¯, one has
= +h¯ +h¯2 + . (3.19)
Leff L2 L4 L6 ···
Here contains all possible contributions with four derivatives, or two derivatives
4
L
and one field χ, or χ2, and similarly for the higher order terms ( contains in
4
L
addition the Wess-Zumino-Witten lagrangian [24]). The generating functional
LWZW
is given by
eih¯Z(v,a,s,p) = [dU]eh¯i Leffdx , (3.20)
Z R
and its low-energy expansion is obtained from Z = Z +h¯Z + . One expands
2 4
···
1
= ¯ +C ξ + ξD ξ +E ξ3 +F ξ4 + ; I = 2,4,... , (3.21)
I I I I I I
L L 2 ···
where ¯ denotesthelagrangian ,evaluatedatthesolutiontotheclassicalequation
I I
L L
of motion δ = 0. (To simplify the notation, we have dropped the SU(2) - indices
2
L
in ξ and inRthe operators CI,DI,... .) The fluctuation ξ is of order h¯1/2. Then one
obtains
eih¯Z = eh¯iScl [dξ]eh¯i 21ξD2ξdx ,
△
Z R
1
= 1 (E ξ3 +h¯C ξ) (E ξ3 +h¯C ξ) dxdy
△ − 2h¯2 2 4 x 2 4 y
Z h i
i
+ 2F ξ4 +h¯ξD ξ dx+O(h¯2) , (3.22)
2 4
2h¯ x
Z h i
with S = dx¯ .
cl Leff
At ordeRr E4, this result amounts to evaluating one-loop graphs generated by
2
L
and adding the tree graphs from +h¯ [4]. These contributions then add up to
2 4
L L
Z , which contains the leading-order term Vµν. It is a specific feature of the process
4 4
γγ π0π0 that the counterterms contained in do not contribute to Vµν–the sum
→ L4 4
of the one-loop graphs is therefore ultraviolet finite [1, 2].
The diagrams which generate Z are displayed in Fig. 2. The solid-dashed lines
6
stand for the propagator D 1, and the framed symbols I denote vertices from
2− LI
according to Eq. (3.21).
In order not to interrupt the argument, we relegate the discussion of the lead-
ing contribution Vµν to appendix B and continue in the following section with the
4
evaluation of the next-to-leading order term Vµν.
6
8
4 Renormalization
TheevaluationofVµν iscomplex. Weoutlineinthisandinthefollowing two sections
6
the procedure – omitting, however, all details.
4.1 The lagrangians and
4 6
L L
The lagrangian contributes to Vµν through one-loop diagrams, see Fig. 2. Its
L4 6
general form is [4]
= (4) + ,
L4 L LWZW
7
(4) = l P + , (4.23)
i i
L ···
i=1
X
where
1
P = uµu 2 ,
1 µ
4h i
1
P = u u uµuν ,
2 µ ν
4h ih i
1
P = χ 2 ,
3 +
16h i
i
P = u χµ ,
4 µ
4h −i
1
P = fµνf ,
5 µν
−2h − − i
i
P = fµν[u ,u ] ,
6 4h + µ ν i
1
P = χ 2. (4.24)
7
−16h −i
Here we used the notation
u = iu D Uu = iuD U u = u ,
µ † µ † − µ † †µ
χ = u χu uχ u ,
† † †
± ±
χµ = u Dµχu uDµχ u ,
† † †
− −
fµν = uFµνu u Fµνu , (4.25)
± L † ± † R
with u2 = U. The quantity Fµν (Fµν) stands for the field strength associated with
R L
the nonabelian external field v +a (v a ).
µ µ µ µ
−
The ellipsis in (4.23) denotes polynomials in the external fields which are inde-
pendent of the pion variables. These do not contribute to S-matrix elements and are
9
therefore not needed in the following. Finally, the anomaly term contributes to
LWZW
Vµν [25]. This is beyond the accuracy of the low-energy expansion considered here.
8
The realization of G on u is
G
u → gRuh† = hugL† , (4.26)
such that
G
I hIh (4.27)
†
→
forthequantitiesin(4.25). Thelow-energyconstantsl aredivergent, exceptl . They
i 7
remove the ultraviolet divergences generated by the one-loopgraphs–we discuss them
in more detail below.
Intheconstructionof (4),theequationofmotionδ = 0hasbeenused. Itcan
2
L L
be shown that adding terms to (4) which vanish uponRuse of the equation of motion
L
affects the generating functional at order E6 by a local term3–these contributions
may thus be omitted.
The lagrangian contributes a polynomial part to Vµν which cancels the ultra-
L6 6
violet singularities generated by the two-loop diagrams. The general structure of
6
L
is not yet available in the literature [26]. Concerning the present calculation, we note
that the lagrangian
1
= f fµσ +f fµσ Tρ + ,
L6 F2h +µρ + −µρ − i σ ···
T = d u u +g d uµu +d χ (4.28)
ρσ 1 ρ σ ρσ 2 µ 3 +
h i { h i h i}
generates a polynomial in A ,B which has the same structure as the divergent part
6 6
in the two-loop contribution,
20
A = [16(d d )M2 +(d +8d )s]+ ,
6 9F4 3− 2 1 2 ···
10
B = d + . (4.29)
6 −9F4 1 ···
We may therefore remove the divergences in Vµν by simply dropping the singular
6
parts in A and B , see below.
6 6
4.2 Regularization and renormalization
We use dimensional regularization and set
ω = d 4 , (4.30)
−
3We thank G. Ecker for an explicit proof of this statement and for illuminating discussions
concerning the material in section 4.2.
10