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Optics Express

Optics Express

  • Editor: C. Martijn de Sterke
  • Vol. 19, Iss. 10 — May. 9, 2011
  • pp: 9714–9736
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Classical and quantum properties of cylindrically polarized states of light

Annemarie Holleczek, Andrea Aiello, Christian Gabriel, Christoph Marquardt, and Gerd Leuchs  »View Author Affiliations


Optics Express, Vol. 19, Issue 10, pp. 9714-9736 (2011)
http://dx.doi.org/10.1364/OE.19.009714


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Abstract

We investigate theoretical properties of beams of light with non-uniform polarization patterns. Specifically, we determine all possible configurations of cylindrically polarized modes (CPMs) of the electromagnetic field, calculate their total angular momentum and highlight the subtleties of their structure. Furthermore, a hybrid spatio-polarization description for such modes is introduced and developed. In particular, two independent Poincaré spheres have been introduced to represent simultaneously the polarization and spatial degree of freedom of CPMs. Possible mode-to-mode transformations accomplishable with the help of Bconventional polarization and spatial phase retarders are shown within this representation. Moreover, the importance of these CPMs in the quantum optics domain due to their classical features is highlighted.

© 2011 OSA

1. Introduction

In this work we aim at establishing a proper theoretical framework for the classical optics description of cylindrically polarized modes (CPMs) of the electro-magnetic field. We cover the subtleties in terms of their description, highlight their symmetry properties and determine their total angular momentum. Finally, we show a suitable way of constructing a reasonable tool of displaying these modes on a defined pair of Poincaré spheres. In particular, in Sec. 1, we illustrate the rotation principles underlying the intrinsic cylindrical polarization of the investigated beams. We introduce all possible configurations of these modes and calculate the Schmidt rank, as this is a convincing way to determine the potential inseparability of these modes. We complete the theoretical investigations by deducing the total angular momentum of these CPMs in Sec. 3. In Sec. 4, we introduce our Poincaré sphere representation for the hybrid polarization/spatial degrees of freedom characterizing CPMs. In Sec. 5, we discuss the manipulations of these states of light on these hybrid Poincaré spheres with the help of conventional phase retarders. Furthermore, in Sec. 6, we extend our formalism to the quantum domain and illustrate the importance of these modes in quantum optics experiments. Finally, in the Appendix, we furnish the reader with the mathematical tools which are essential to fully cover the topics of the main text.

2. Classical description of light beams with cylindrical polarization

2.1. Introduction to rotation principles

To mathematically introduce the radially and azimuthally polarized vector fields, let us consider a well collimated, monochromatic light beam with the wavelength λ = 2π/k propagating along the z-axis. The corresponding electric field can be written as
E(r,t)=12(u(r)eiχ+u*(r)eiχ),
(1)
where χ = k(ctz) – π/2 [14

14. R. Loudon, The Quantum Theory of Light (Oxford University Press, 2000).

] and r = x + y ŷ + z . Furthermore, u(r) is a paraxial mode function of the form
u(r)=f1(r)x^+f2(r)y^.
(2)

The unit vectors and ŷ indicate the polarization in the x and y directions and f 1(r) and f 2(r) are two independent solutions of the paraxial wave equation.:
2f1,2x2+2f1,2y2+2ikf1,2z=0.
(3)

Equation (14) has some interesting consequences: First, we note that, since x = r cos θ and y = r sin θ, we can rewrite x = x(r,θ) and y = y(r,θ), having assumed r,θ as independent variables. Then, by definition,
x(r,θ±φ)=rcos(θ±φ)=r(cosθcosφsinθsinφ),
(16)
and
y(r,θ±φ)=rsin(θ±φ)=r(sinθcosφ±cosθsinφ).
(17)
These two equations can be rewritten in compact vector notation as:
x(r,θ±φ)=R(±φ)x(r,θ).
(18)
Now, since the argument of the functions u ± (r,θ) = u ± (x) on the left side of Eq. (14) is exactly x(r,θ ± φ), we can use the result above to rewrite Eq. (14) as:
u±(R(±φ)x)=R(φ)u±(x),
(19)
or, equivalently,
u+(x)=R(φ)u+(R(φ)x),
(20a)
u(x)=R(φ)u(R(φ)x).
(20b)
At this point, one should remember that any 2-dimensional vector field V(x,y) transforms under a global counterclockwise active rotation by an angle φ into the new field W(x,y) whose functional expression is determined according to the following rule:
W(x)=R(φ)V(R(φ)x).
(21)
Thus Eq. (20a) coincides exactly with the definition of a vector field which is invariant with respect to a global rotation by an angle φ. This means that co-rotating beams have a polarization pattern invariant with respect to global rotations. This result is intuitively obvious for radial and azimuthally polarized beams. Conversely, the corresponding Eq. (20b) for the counter-rotating beams has a less straightforward geometric interpretation. For this case, one can show that Eq. (20b) can be recast in the form of a global rotation as:
R(2φ)u(x)=R(φ)u(R(φ)x).
(22)
This equation simply shows that a global rotation by an angle φ upon the counter-rotating modes is equivalent to a local rotation by an angle 2φ on the same mode.

2.2. Deduction of cylindrically polarized modes

As we are now knowledgeable about all rotation properties which the cylindrically polarized optical beams have to obey to, we want to explicitly derive all possible cylindrical configurations of these mode functions which fulfill Eq. (14) taken with the “+” sign. However, it will be eventually shown that these coincide with the well-known radially and azimuthally polarized beams.

After inverting Eq. (5) and inserting the result into Eq. (2), we obtain as a paraxial mode function
u(r,θ)=r^(cosθf1+sinθf2)+θ^(sinθf1+cosθf2).
(23)
By comparing Eq. (23) with Eq. (13), one obtains at once:
f1(r)=a(r)cosθb(r)sinθ,
(24a)
f2(r)=a(r)sinθ+b(r)cosθ.
(24b)
Since x = r cos θ and y = r sin θ, Eq. (24a) and Eq. (24b) can be rewritten as:
f1(r)=xα(r)yβ(r),
(25a)
f2(r)=yα(r)+xβ(r),
(25b)
where α(r) and β(r) can be determined by keeping in mind that f 1 and f 2 need to satisfy the paraxial Eq. (3). In particular, we want to express the functions f 1(r) and f 2(r) in terms of the first-order Hermite-Gauss [15

15. M. J. Padgett and J. Courtial, “Poincaré-sphere equivalent for light beams containing orbital angular momentum,” Opt. Lett. 24(7), 430–432 (1999). [CrossRef]

] solutions of the paraxial wave equation which can be – apart from an irrelevant normalization factor 8/π – written at z = 0, as
ψ10(x)=xe(x2+y2)/w02/w0,
(26a)
ψ01(x)=ye(x2+y2)/w02/w0,
(26b)
where w 0 is the beam waist. Thus, from Eq. (25a) and Eq. (25b) and the fact that xψ 10 and yψ 01, it follows that:
f1(r)=12(Aψ10Bψ01),
(27a)
f2(r)=12(Bψ10+Aψ01),
(27b)
where A, B are arbitrary numerical constants. In general, with ψnm = ψnm(x,y,z) (n,m ∈ ℕ0) we denote the Hermite-Gauss solutions of the paraxial wave equation of the order N = n + m [15

15. M. J. Padgett and J. Courtial, “Poincaré-sphere equivalent for light beams containing orbital angular momentum,” Opt. Lett. 24(7), 430–432 (1999). [CrossRef]

]. Consequently, all possible solutions which fulfill the global law of Eq. (14) with the “+” sign, obviously have the form of:
u=12(Aψ10Bψ01)x^+12(Bψ10+Aψ01)y^.
(28)
By choosing either {A = 1,B = 0} or {A = 0,B = 1} and inserting this in Eq. (28), one obtains after normalization:
u^R+=12(ψ10x^+ψ01y^),
(29a)
u^A+=12(ψ01x^+ψ10y^).
(29b)
The same steps that lead us to Eq. (29a) and Eq. (29b) could be repeated for the counter-rotating modes. However, it is more instructive to note that, from a visual inspection, one can immediately deduce that u^R+ and u^A+ live in a four dimensional space spanned by the basis formed by the Cartesian product of the {ψ 10,ψ 01} mode bases and the polarization vectors {,ŷ}:
{ψ10,ψ01}{x^,y^}={ψ10x^,ψ10y^,ψ01x^,ψ01y^}.
(30)
However, the two vectors { uR+,uA+} only form a two dimensional basis and consequently two more vectors have to be added to span the whole space. By applying a unitary transformation to such a basis, we can obtain a new set of basis vectors, namely { uR,uA}, where:
u^R=12(ψ10x^+ψ01y^),
(31a)
u^A=12(ψ01x^+ψ10y^).
(31b)
The vectors { u^R+,u^A+,u^R,u^A}, whose intensity and polarization profiles are illustrated in Fig. 2, form a complete four-dimensional basis, orthonormal with respect to the scalar product defined as:
Fig. 2 Complex polarization patterns of (a) uR+, (b) uA+, (c) uR, (d) uA, underlayed with the doughnut shaped intensity distribution.
(u(x),v(x))=i=12ui*(x)vi(x)dxdy,
(32)
where u(x) and v(x) are two arbitrary vector functions. It should be stressed that the basis vectors u^R and u^A do fulfill the rotation law of Eq. (14), but with the “−” sign:
u(r,θα)=R(α)u(r,θ).
(33)
Furthermore, Eq. (14) and Eq. (33) imply that these two sets of modes do not mix under rotation. This is interesting, as many recent experiments show [6

6. R. Dorn, S. Quabis, and G. Leuchs, “Sharper focus for a radially polarized light beam,” Phys. Rev. Lett. 91(23), 233901 (2003). [CrossRef] [PubMed]

, 12

12. C. Gabriel, A. Aiello, W. Zhong, T. G. Euser, N. Y. Joly, P. Banzer, M. Förtsch, D. Elser, U. L. Andersen, Ch. Marquardt, P. St. J. Russell, and G. Leuchs, “Entangling different degrees of freedom by quadrature squeezing cylindrically polarized modes,” Phys. Rev. Lett. 106(6), 060502 (2011). [CrossRef] [PubMed]

] that co- and counter-rotating modes do not occur simultaneously.

2.3. Schmidt rank of radially and azimuthally polarized modes

Another important property of CPMs can be characterized by the so-called Schmidt rank of the modes. The Schmidt rank [16

16. A. Peres, Quantum Theory: Concept and Methods (Kluwer Academic, Boston, 1995).

] is an intriguing quantity as it provides information about a potential inseparability of these modes. In the following, we demonstrate how to determine this rank for the radially and azimuthally polarized optical beams. This can be similarly calculated for the counter-rotating modes.

If one has a look at the structure of these basis vectors, one can easily see that they are clearly non-separable, in the sense that it is not possible to write any of these vector functions as the product of a uniformly polarized vector field times a scalar function. Such a “degree of non-separability” may be quantified by the Schmidt-rank of the function, remembering that all separable functions have a Schmidt rank of 1. Conversely, we are now going to prove that u^R+ and u^A+ admit a Schmidt decomposition with rank 2:

Let ê 1 and ê 2 be two orthogonal unit vectors defined as ê 1 = and ê 2 = ŷ, with
ei*·ej=δij,i,j{1,2},
and let v 1 and v 2 be two orthogonal functions defined as
v1=ψ10(x,y,z),v2=ψ01(x,y,z),
(34)
where, because of the orthogonality of the Hermite-Gauss modes, one has:
vi*vjdxdy=δij,i,j{1,2}.
Then we can immediately write down, for example for the u^R+ mode:
u^R+=i=12(λi)1/2e^ivi,
(35)
with λ 1 = λ 2 = 1/2. The right side of Eq. (35) represents the Schmidt decomposition of the vector function u^R+ with rank
K=1/i=12λi2=2.
(36)

For the function u^A+ the same demonstration holds but with v 1 and v 2 defined as
v1=ψ01(x,y,z),v2=ψ10(x,y,z),
(37)
This completes the proof.

3. Angular momentum of cylindrically polarized optical beams

To complete our discussion about the properties of the “+” and “−” vector bases, we note that all four vector functions { uR+,uR,uA+,uA} have the total (orbital plus spin) angular momentum (TAM) of zero which is deduced below. Following Berry [17

17. M. V. Berry, “Optical currents,” J. Opt. A, Pure Appl. Opt. 11(9), 094001 (2009). [CrossRef]

], we can write the “local” (position-dependent) density of the linear momentum carried by the paraxial mode u of the electromagnetic field as:
p(r)=Im[u*·()u]+12Im[×(u*×u)]porb(r)+psp(r),
(41)
where ∇ = x + ŷy and u*·()u=iui*(ui). A prefactor of c 2 ͛ 0/ω has been omitted. The symbol “×” denotes the ordinary cross product in ℝ3. In Eq. (41), the first term p orb(r) is called the orbital Poynting current and is independent of the polarization state of the beam. Conversely the second term, the so-called spin Poynting current p sp(r), contains the local “density of spin” of the beam via the term Im (u* × u) [17

17. M. V. Berry, “Optical currents,” J. Opt. A, Pure Appl. Opt. 11(9), 094001 (2009). [CrossRef]

]. An explicit calculation from Eq. (41) furnishes:
px(r)=1kIm[f1*(xf1)+f2*(xf2)]+1kIm[y(f1*f2)],
(42a)
py(r)=1kIm[f1*(yf1)+f1*(yf2)]1kIm[xf1*f2],
(42b)
pz(r)=|f1|2+|f2|2,
(42c)
where k indicates the wave vector k = ω/c. The power of a light beam as actually measured by a photo-detector in the laboratory equals the flow of the Poynting vector of the beam across the detector surface. Since the electromagnetic linear momentum density p is equal to 1/c 2 times the Poynting vector, it follows that a quantity experimentally observable is P·, where
P=p(r)dxdyPorb+Psp,
(43)
is the linear momentum of the field per unit length. Without loss of generality, we have assumed that the detector surface coincides with the xy-plane normal to the beam propagation axis . The double integration in Eq. (43) is extended to the whole xy-plane. This means that for any pair of functions f 1 and f 2, normalizable in the xy-plane, which drop to zero at infinity, the “spin” surface terms y(f1*f2) and x(f1*f2) on the right sides of Eq. (42a) and Eq. (42b) respectively, do not contribute to the measured momentum. In other words one has P sp = 0. It is worth noting that although f 1 and f 2 are both functions of r, as opposed to x we can practically evaluate Eq. (43) at z = 0 because the quantity P is independent of z [18

18. A. Aiello, N. Lindlein, Ch. Marquardt, and G. Leuchs, “Transverse angular momentum and geometric spin hall effect of light,” Phys. Rev. Lett. 103(10), 100401 (2009). [CrossRef] [PubMed]

].

The total angular momentum density is
j(r)=r×p(r)=r×[porb(r)+psp(r)]l(r)+s(r),
(44)
where the orbital angular momentum density is
lx(r)=y(|f1|2+|f2|2)zkIm(f1*yf1+f2*yf2),
(45a)
ly(r)=x(|f1|2+|f2|2)+zkIm(f1*xf1+f2*xf2),
(45b)
lz(r)=Im[x(f1*yf1+f2*yf2)y(f1*xf1+f2*xf2)],
(45c)
and the spin angular momentum density is
sx(r)=zkIm[x(f1*f2)],
(46a)
sy(r)=zkIm[y(f1*f2)],
(46b)
sz(r)=1kIm[xx(f1*f2)+yy(f1*f2)].
(46c)
We proceed as in the linear momentum case and write the angular momentum of the field per unit length as
J=j(r)dxdyL+S,
(47)
were, similarly to P, also L and S are independent of z [18

18. A. Aiello, N. Lindlein, Ch. Marquardt, and G. Leuchs, “Transverse angular momentum and geometric spin hall effect of light,” Phys. Rev. Lett. 103(10), 100401 (2009). [CrossRef] [PubMed]

]. Again, for any pair of functions f 1(r) and f 2(r) that vanish sufficiently fast for |xy| → ∞, integration by parts leads to Sx = 0 = Sy, and to
[xx(f1*f2)+yy(f1*f2)]dxdy=2f1*f2dxdy.
(48)
From this relation it follows straightforwardly that
Sz=ik(f1*f2f1f2*)dxdy.
(49)
Equation (49) clearly reproduces the well-know expression of the helicity Sz = i(αβ* −α*β)/k of a uniformly-polarized beam when one chooses f 1 = αψ(r) and f 2 = βψ(r), where ψ(r) is a normalizable function in the xy-plane and α, β a pair of complex number such that |α|2 +|β|2 = 1.

At this point, we can easily evaluate J for the set of our four cylindrically polarized modes without performing any actual calculations. In fact, since J is independent of z, we can evaluate it at z = 0 where both ψ 10(x,y,z = 0) and ψ 01(x,y,z = 0) are real-valued functions. Then from Eqs. (45–46) it follows at once that S = 0 and that
Lx=y(|f1|2+|f2|2)dxdy,
(50a)
Ly=x(|f1|2+|f2|2)dxdy,
(50b)
Lz=0.
(50c)
However, it is easy to see from Eq. (26a) and Eq. (26b) that by definition |ψ 10(x,y,0)|2 and |ψ 01(x,y,0)|2 are even functions with respect to both x and y coordinates. This automatically implies, because of symmetry, that Lx = 0 = Ly.

We have thus demonstrated that the TAM of our modes is strictly zero. Alternatively, the same result could have been found in a more intuitive but less formal way by writing the modes { uR+,uR,uA+,uA} in the left/right circular polarization basis and in the ϕ±1,0LG Laguerre-Gauss spatial basis. In this case, both the polarization and spatial modes carry a unit of angular momentum which may be summed or subtracted from each other, thus leading to a straightforward physical interpretation.

4. Hybrid Poincaré sphere representation

In general, fully polarized optical beams, similarly to quantum pure states of two-level systems [19

19. J. M. Jauch and F. Rohrlich, The Theory of Photons and Electrons: The relativistic Field Theory of Charged Particles with Spin one-half (Addison-Wesley Publishing Company, 1955).

], can be described by points on a sphere of unit radius [20

20. R. M. A. Azzam and N. Bashra, Ellipsometry and Polarized Light (North Holland Personal Library, 1988).

], the so-called polarization Poincaré sphere [PPS, Fig. 4(b)]. The North and South Poles of this sphere represent left-and right-handed polarization states that correspond to ±1 spin angular momentum eigenstates, respectively. Similarly, the spatial-mode Poincaré sphere [SMPS, Fig. 4(a)] [15

15. M. J. Padgett and J. Courtial, “Poincaré-sphere equivalent for light beams containing orbital angular momentum,” Opt. Lett. 24(7), 430–432 (1999). [CrossRef]

] sticks to the same superposition principle but in terms of the spatial distribution of the optical beams. In this case, the North and the South Pole of the sphere represent optical beams in the Laguerre-Gaussian modes ϕ±1,0LG, which correspond to ±1 orbital angular momentum eigenstates [21

21. L. Allen, M. W. Beijersbergen, R. J. C. Spreeuw, and J. P. Woerdman, “Orbital angular momentum of light and the transformation of laguerre-gaussian laser modes,” Phys. Rev. A 45(11), 8185–8189 (1992). [CrossRef] [PubMed]

], respectively. Each point on the Poincaré sphere, both on the polarization and the spatial-mode one, may be put in one-to-one correspondence with a unit three-dimensional real vector S⃗ ={S 1/S 0,S 2/S 0,S 3/S 0}, where S02=S12+S22+S32, as shown in Fig. 3. The three components Si, i ∈ {1,2,3} are known as the Stokes parameters [20

20. R. M. A. Azzam and N. Bashra, Ellipsometry and Polarized Light (North Holland Personal Library, 1988).

] and S 0 denotes the total intensity of the beam.

Fig. 3 Poincaré sphere representation for an arbitrary two-dimensional system. Here, the radius r of the sphere is fixed to one, and spherical angles {ϑ,ϕ} are related to the Stokes parameters via the relation: {cos ϑ,sin ϑ cos ϕ,sin ϑ sin ϕ} = {S 1/S 0, S 2/S 0, S 3/S 0}, where S 0 denotes the total intensity of the beam.
Fig. 4 Fundamental idea behind the new Poincaré sphere representation: Combining the SMPS (a) and the PPS (b) to the hybrid Poincaré spheres (HPSs) (c) and (d).

In this section, we present a novel representation for cylindrically polarized optical beams based on the simultaneous use of both the PPS and the SMPS representation. We show that it is possible, under certain assumptions, to connect the two polarization and spatial representations to obtain two hybrid Poincaré spheres (HPSs) that embody all characteristics of these complex modes of the electromagnetic field. Up to now, only uniform polarization patterns could have been displayed as points on the PPS. However, with this new representation, nonuniform polarization patterns are represented for the first time. This is especially helpful – let it be in theoretical considerations or in experiments – if one investigates effects deriving from these spatially dependent, complex polarization patterns. Furthermore, manipulations of states within this representation in accordance with the conventional tools such as the PPS and SMPS can be carried out. This is investigated in Sec. 5.

In principle, the two-dimensional polarization and spatial spaces, respectively spanned by the bases {,ŷ} and {ψ 10,ψ 01}, may be combined in several different manners to produce a four-dimensional mixed polarization-spatial mode space. A complete representation of such a space would require the simultaneous display of eight real numbers which is impossible in a three dimensional space. Consequently, our aim is to find an appropriate way for describing this space by using conventional tools, such as the PPS or the SMPS respectively. We begin by noticing that if we are permitted to discard some information, then a lower-dimension visual representation may be possible. The characteristics of this picture would then be fixed by the kind of degree of freedom (DOF) we choose to represent. Driven by certain recent experimental results [12

12. C. Gabriel, A. Aiello, W. Zhong, T. G. Euser, N. Y. Joly, P. Banzer, M. Förtsch, D. Elser, U. L. Andersen, Ch. Marquardt, P. St. J. Russell, and G. Leuchs, “Entangling different degrees of freedom by quadrature squeezing cylindrically polarized modes,” Phys. Rev. Lett. 106(6), 060502 (2011). [CrossRef] [PubMed]

], here we look for a hybrid representation, which would be suitable to describe physical situations where co- and counter-rotating modes do not mix with each other. This absence of mixing occurs, for example, when the light propagates through cylindrically symmetric optical systems as fibres, non-astigmatic lenses, et cetera. In this case, the disregarded information would amount to the relative phase and amplitude between the sets of co- and counterrotating modes. The representation deriving from these constraints is illustrated in Fig. 4. The North and South Pole of sphere (c) represent radially and azimuthally polarized states whose total angular momentum (orbital plus spin) is equal to zero. Analogously, sphere (d) represent counter-radial and counter-azimuthal states, still with total zero angular momentum. Each of these two pairs of states represent a novel kind of hybrid DOF of the electromagnetic field, since such states describe neither purely polarized nor purely spatial modes of the field.

Mathematically, the combination of the SMPS and the PPS is realized by forming the four dimensional basis given by the Cartesian product of the the spatial mode {ψ 10,ψ 01} and polarization {,ŷ} bases:
{ψ10,ψ01}{x^,y^}={ψ10x^,ψ10y^,ψ01x^,ψ01y^}.
(51)
To span this four dimensional space, we choose the basis { u^R+,u^A+,u^R,u^A}. As it was shown above, these two sets of vector functions, namely the “+” and “−” ones, do not mix under rotation and, moreover, obey the two fundamental rotation laws in Eq. (14) and Eq. (33). Such a physical constraint may be mathematically implemented by a superselection rule [16

16. A. Peres, Quantum Theory: Concept and Methods (Kluwer Academic, Boston, 1995).

] which forbids interference between states separated according to the law in Eq. (14). Therefore, we use this law to split the four dimensional space {ψ 10 ,ψ 10 ŷ,ψ 01 ,ψ 01 ŷ} into the Cartesian sum of two subspaces spanned by the “+” and “−” sets of modes:
{ψ10,ψ01}{x^,y^}={uR+,uA+}{uR,uA}.
(52)
This shows, that depending upon their behavior under rotation, cylindrically polarized optical beams may be subdivided in two independent sets which can be represented by points on the surface of two distinct hybrid Poincaré spheres, as illustrated in Fig. 4. It is worth to stress once again that a beam represented by an arbitrary superposition of either the standard basis vectors {ψ 10 ,ψ 10 ŷ,ψ 01 ,ψ 01 ŷ} or the cylindrical basis vectors { uR+,uA+,uR,uA} needs four complex numbers to be described completely. This amounts to eight real numbers which can be reduced to seven because of normalization. However, our superselection rule defined above reduces the number of real parameters necessary to describe cylindrically polarized beams further to 3 ⊕ 3, thus permitting the introduction of a “two-Poincaré Cartesian sum” representation.

Besides evident visualization properties, our representation permits to describe straight forwardly cylindrically polarized states of light, in terms of hybrid Stokes parameters, as opposed to the either polarization or spatial Stokes parameters. These hybrid Stokes parameters convey information about both polarization and spatial-mode degrees of freedom simultaneously. They are naturally defined as:
S0±=fR±*fR±+fA±*fA±,
(53a)
S1±=fR±*fR±fA±*fA±,
(53b)
S2±=fR±*fA±+fA±*fR±,
(53c)
S3±=i(fR±*fA±fA±*fR±),
(53d)
where the symbol * denotes the complex conjugation. fA±=(uA±,E) and fR±=(uR±,E) are the field amplitudes of the electric field vector in the bases { uA±,uR± }, where
E=fA±uA±+fR±uR±.
(54)
These amplitudes can be expressed in terms of the spherical coordinates on the two independent hybrid Poincaré spheres (HPSs) as
fA±=cos(θ/2),
(55a)
fR±=exp(iϕ)sin(θ/2).
(55b)
The first sphere of the HPSs represents the “+” modes [Fig. 5(a)] and the second the “−” [Fig. 5(b)] modes. It is possible to show that these Stokes parameters can be actually measured by means of conventional optical elements. To perform such a measurement on either of the HPSs one needs to project an arbitrary cylindrically polarized state of light onto the Stokes basis modes of interest. This can be achieved by a combination of spatial optical elements, such as asymmetric Mach Zehnder interferometers [22

22. H. Sasada and M. Okamoto, “Transverse-mode beam splitter of a light beam and its application to quantum cryptography,” Phys. Rev. A 68(1), 012323 (2003). [CrossRef]

, 23

23. M. T. L. Hsu, W. P. Bowen, and P. K. Lam, “Spatial-state stokes-operator squeezing and entanglement for optical beams,” Phys. Rev. A 79(4), 043825 (2009). [CrossRef]

], and polarization optical elements, such as waveplates and polarization beam splitters (PBSs). A scheme to measure the S1+, S2+ or S3+ basis modes is illustrated in Fig. 6. Being able to do so is particularly relevant for the possible quantum applications of our formalism, where simultaneous measurability of Stokes parameters describing spatially separated optical beams, is crucial [24

24. N. Korolkova, G. Leuchs, R. Loudon, T. C. Ralph, and Ch. Silberhorn, “Polarization squeezing and continuous-variable polarization entanglement,” Phys. Rev. A 65(5), 052306 (2002). [CrossRef]

]. Concerning the structure of the HPSs, an interesting feature is that every point on the meridian between the North and the South Pole (θ ± ∈ [0,π], ϕ ± = 0) describes a locally linear polarization state as illustrated in both Fig. 5(a) and Fig. 5(b). Except for these points and the two ones on the S 3 axis ( θ=π2,ϕ={π2,32π}) which represent circularly polarized states, all remaining points on the spheres describe states of non-uniform elliptical polarization. This is fully consistent with the structure of the conventional PPS or SMPS representation.

Fig. 5 Polarization states on the HPSs, represented on the sphere of the “+” modes (a) and the “−” modes (b). The Stokes parameters on both the “+” and “−” hybrid Poincaré sphere are explicitly marked.
Fig. 6 (a) Experimental setup to measure hybrid Stokes parameters on the “+” HPS: Measurement of S1+: (A) = (B) = (C) = (D) no optical element. Measurement of S2+: (A) = (C) = (D) HWP 22.5°, (B) HWP 112.5°. Measurement of S3+: (A) = (C) HWP 22. 5°, QWP 0°, (B) HWP 112.5°, QWP 90°, (D) HWP 22.5°, QWP 90°. S0+ can be measured in every case if one uses the sum signal between both ports. In (b), the asymmetric Mach Zehnder interferometer is displayed which acts as a spatial mode splitter (SMS) and has been used three times is the Stokes measurement setup.

5. Manipulating states on the HPSs

As in terms of the conventional PPS and SMPS representation, conversions of states by means of polarization and spatial phase retarders such as waveplates or cylindrical lenses, respectively, are possible. In general, waveplates only act on the polarization part of the vector beam and cylindrical lenses on the spatial mode part, as it was shown in [21

21. L. Allen, M. W. Beijersbergen, R. J. C. Spreeuw, and J. P. Woerdman, “Orbital angular momentum of light and the transformation of laguerre-gaussian laser modes,” Phys. Rev. A 45(11), 8185–8189 (1992). [CrossRef] [PubMed]

]. In terms of the PPS, two independently adjustable quaterwaveplates (QWPs) and one halfwaveplate (HWP) are sufficient to control polarization [20

20. R. M. A. Azzam and N. Bashra, Ellipsometry and Polarized Light (North Holland Personal Library, 1988).

], so reaching every point on this sphere is possible. In analogy to this, manipulating states on the SMPS is possible, as well. However, one cannot use wave-plates but instead cylindrical lenses [15

15. M. J. Padgett and J. Courtial, “Poincaré-sphere equivalent for light beams containing orbital angular momentum,” Opt. Lett. 24(7), 430–432 (1999). [CrossRef]

] to control the spatial mode. Displaying the information from both the PPS and the SMPS simultaneously with the new representation, combinations of both waveplates and cylindrical lenses are necessary. The process of converting a state on the HPSs to a different one – for instance, going from S1± via S2±, S1± and S2± back to S1± – can be achieved by applying two halfwaveplates and rotating one of these as illustrated in Fig. 7. The fixed HWP induces a local flip of the vectors in the polarization pattern and the other one rotates continuously these vectors. The equivalent effect can also be achieved by means of spatial optical elements. For example, one could exploit two pairs of cylindrical lenses to perform the rotations described before. This is also illustrated in Fig. 7. It has to be noted that most of the possible mode-to-mode transformations can be carried out with either polarization or spatial phase retarders. This is due to the structural inseparability of these two degrees of freedom in the cylindrically polarized modes which has been shown earlier. A pair of cylindrical lenses acts on the pair of spatial modes {ψ 10,ψ 01} exactly in the same way as a HWP acts on the pair {,ŷ} in the polarization space [21

21. L. Allen, M. W. Beijersbergen, R. J. C. Spreeuw, and J. P. Woerdman, “Orbital angular momentum of light and the transformation of laguerre-gaussian laser modes,” Phys. Rev. A 45(11), 8185–8189 (1992). [CrossRef] [PubMed]

]. This principle becomes clear if one looks on the transformation of a radially polarized light beam to an azimuthally polarized one which can be achieved by means of either spatial or polarization optics:
Fig. 7 Manipulation of states on the HPSs, represented on the sphere of the “+” modes (a) and the “−” modes (b). This special case can be carried out by both waveplates and pairs of cylindrical lenses. The distance between the (c) two cylindrical lenses is 2f, where f denotes their focal length.
u^R+x^ψ10+y^ψ01fixedspatialmodeHWP,HWPu^A+y^ψ10+x^ψ01,u^R+x^ψ10+y^ψ01fixedpolarizationtwopairsofcyl.lensesu^A+x^ψ01+y^ψ10.
However, due to the special cylindrical symmetry of our states, no “arbitrary-to-arbitrary” [25

25. J. N. Damask, Polarization Optics in Telecommunication (Springer, 2005).

] transformations with the help of waveplates or cylindrical lenses are permitted. Therefore, only a few transformations can be accomplished without altering the symmetry of the state. These are listed below:
{θ,ϕ}{θ=πθ,ϕ}ϕ{0,π2,π,3π2},
(56a)
{θ,ϕ}{θ=πθ,ϕ=π+ϕ}ϕ{0,π2,π,3π2},
(56b)
{θ,ϕ}{θ=θ,ϕ=π+ϕ}ϕ[0,2π[,
(56c)
where θ ∈ [0,π] and ϕ refers to the angles parametrizing a sphere, as shown in Fig. 8. These constraints are due to the fact that phase retarders imprint a spatially uniform phase shift onto the cross section of the beam, which is not, per definition, cylindrically invariant. These constraints are explicitly deduced in the Appendix.

Fig. 8 Transformations on the “+” HPS: Without altering the rotational symmetry of the state, these transformations are possible with the help of conventional phase retarders.

To complete the discussion about possible manipulations, we notice that the two sets { uR+,uA+}and { uR,uA} are connected to each other by a mirror image symmetry operation, which is illustrated in Fig. 9. This can be practically performed by a halfwaveplate [25

25. J. N. Damask, Polarization Optics in Telecommunication (Springer, 2005).

]. So, jumps between the “+” and the “−” spheres can be carried out easily and they considerably enlarge the number of possible transformation.

Fig. 9 By a mirror image symmetry operation one can pass from one sphere to the respective other sphere. This action is practically performed by means of a HWP.

6. Quantum properties of cylindrically polarized states of light

As we have discussed only classical properties of cylindrically polarized states of light so far, we cover their quantum aspects in the following. In this section, we give a few examples of application of our hybrid polarization/spatial formalism to some quantum states of light.

To begin with, let us write the electric-field quantum operator for paraxial beams of light in the same units of Eq. (1) as:
E^(r,t)=E^+(r,t)+E^(r,t),
(57)
where Ê = (Ê +), and
E^+(r,t)=12λ=12n,me^λa^λnmψnmeiχ,
(58)
with ê 1 = and ê 2 = ŷ and n,m ∈ {0,1,2, …}. The paraxial annihilation operators âλnm satisfy the canonical commutation relations
[a^λnm,a^λnm]=δλλδnnδmm.
(59)

In order to have a correct description of quantum noise in light beams with cylindrical polarization, we need to establish first a correspondence between the classical and the quantum representation of these beams. For sake of simplicity, here we restrict our attention to co-rotating modes solely, namely uA+ and uR+. Having in mind this goal, let u AB(r) be the generic cylindrical mode (the superscript “+” will be omitted from now on) that now we write as
uAB(r)=12[x^(Aψ10Bψ01)+y^(Bψ10+Aψ01)]=AuR+BuA,
(60)
where Eq. (29a) and Eq. (29b) were used in the second line and again |A|2 +|B|2 = 1 is assumed. Since (u R, u A) = 0, then two different modes u AB(r) and u A B (r) satisfy the relation
(uAB(r),uAB(r))=A*A+B*B.
(61)
Now the question is: What is the correct form of the field operator âAB that annihilates a photon in the mode u AB(r)? We seek an answer to this question by requiring that:
  1. The coherent state |α=exp(αa^ABα*a^AB)|0 must produce a coherent signal 𝒮 = 〈α|Ê|α〉 equal to the classical electric field:
    𝒮=12[αuAB(r)eiχ+α*uAB*(r)eiχ].
    (62)
  2. The single-photon state |1=a^AB|0 must generate a photon wave-function 〈0|Ê +|1〉 equal to:
    0|E^+|1=12uAB(r)eiχ.
    (63)
With the help of Eq. (60) we guess the following form for the sought annihilation operator âAB:
a^AB=A(a^x10+a^y012)+B(a^x01+a^y102),
(64)
where we have used the polarization indexes x for λ = 1 and y for λ = 2. From Eqs. (59,64) it follows that:
[a^AB,a^AB]=A*A+B*B,
(65)
which, together with Eq. (61), shows that âA B and a^AB commute whenever the two modes u AB(r) and u A B (r) are orthogonal.

6.1. Coherent states

Let i(β), i ∈ {1,2, … ,∞} be the coherent-state displacement operator defined as
D^i(β)=exp(βa^iβ*a^i),
(66)
where the single label i embodies λ,n,m, so that Σi = Σλ Σn Σm, and we choose the first 4 values of the index i in such a way that
a^1=a^x10,a^2=a^y01,a^3=a^x01,a^4=a^y10.
(67)
If we rewrite the field mode functions as
vi(r,t)e^λψnmexp(iχ),
(68)
then Eq. (58) takes the simpler form
E^+=12i=1via^i.
(69)
Since any coherent state |βi = i(β)|0〉 satisfies the relation
E^+|βi=12j=1vja^j|βi=12viβ|βi,
(70)
it immediately follows that if we define |α 1,α 2,α 3,α 4〉 ≡ 4(α 4) 33) 2(α 2) 1(α 1)|0〉, then
E^+|α1,α2,α3,α4=12i=1via^i|α1,α2,α3,α4=12i=14viαi|α1,α2,α3,α4,
(71)
which trivially implies that
𝒮=α1,α2,α3,α4|E^++E^|α1,α2,α3,α4=12i=14(viαi+vi*αi*)=12[x^(α1ψ10+α3ψ01)+y^(α4ψ10+α2ψ01)]eiχ+c.c.
(72)
Comparison of Eq. (72) with Eq. (60) shows that if we choose
α1=α2=A*α2,α3=α4=B*α2,
(73)
then we have:
D^4(α4)D^3(α3)D^2(α2)D^1(α1)=expi=14(αia^i+αi*a^i)=exp(αa^ABα*a^AB),
(74)
and Eq. (62) is automatically satisfied.

6.2. Fock states

The validity of Eq. (63) can be checked by defining the single-photon ladder operator Â(ξ) as:
A^(ξ)=i=14ξi*a^i,withi=14|ξi|2=1,
(75)
that satisfies the canonical commutation relations
[A^(ξ),A^(ξ)]=1.
Let |1〉 be the single-photon state defined as |1〉 = Â (ξ)|0〉. Then
0|E^+|1=120|i=1via^ij=14ξja^j|0=12i,jξjvi0|a^ia^j|0=12i=14viξi.
(76)
Now, if we choose
ξ1=ξ2=A*12,ξ3=ξ4=B*12,
(77)
then we obtain
A^(ξ)=A(a^x10+a^y012)+B(a^x01+a^y102)=a^AB,
(78)
in agreement with Eq. (63). Thus, we have proved that Eq. (64) furnishes the correct expression for âAB.

6.3. Squeezed states

As a last example, here we write explicitly the expression for an azimuthally polarized squeezed state. The extension to the other cylindrically polarized states is straightforward and, therefore, will be omitted. As a result of such calculation, we shall find that classical inseparability automatically lead to quantum entanglement when the azimuthally polarized beam is prepared in a nonclassical state [12

12. C. Gabriel, A. Aiello, W. Zhong, T. G. Euser, N. Y. Joly, P. Banzer, M. Förtsch, D. Elser, U. L. Andersen, Ch. Marquardt, P. St. J. Russell, and G. Leuchs, “Entangling different degrees of freedom by quadrature squeezing cylindrically polarized modes,” Phys. Rev. Lett. 106(6), 060502 (2011). [CrossRef] [PubMed]

].

The “natural” definition of squeezing operator for the azimuthally polarized field is
S^A(ζ)=exp[12ζ*(a^A)212ζ(a^A)2],
(79)
where ζ= se is the complex squeezing parameter. By substituting Eq. (64) evaluated for A = 0, B = 1 into Eq. (79), it is not difficult to obtain
S^A(ζ)=S^3(ζ/2)S^4(ζ/2)S^34(ζ/2),
(80)
where we have defined the one- and two-mode squeezing operator
S^i(ζ)=exp[12ζ*(a^i)212ζ(a^i)2],
(81)
S^i,j(ζ)=exp(ζ*a^ia^jζa^ia^j).
(82)
Thus, the squeezed-vacuum state
|ζA=S^3(ζ/2)S^4(ζ/2)S^34(ζ)|0.
(83)
is clearly entangled, since
S^34(ζ)|0=1coshsn=0(eiϑtanhs)n|n3|n4,
(84)
represents the well-known entangled two-mode squeezed vacuum [26

26. W. P. Bowen, R. Schnabel, P. K. Lam, and T. C. Ralph, “Experimental investigation of criteria for continuous variable entanglement,” Phys. Rev. Lett. 90(4), 043601 (2003). [CrossRef] [PubMed]

]. Thus, we have shown that classical inseparability naturally lead to quantum entanglement.

7. Conclusion

In conclusion, we have established a proper theoretical framework for a classical description of cylindrically polarized states of the electromagnetic field. We have highlighted the rotation principles of these modes which can be derived from their very particular intrinsic symmetries. Moreover, we found that all CPMs have a zero total angular momentum and are clearly non-separable which can be proven by their Schmidt rank. This subtlety leads to intriguing features of these modes in both the classical and quantum domain of optics, the latter being recently demonstrated in experiments [12

12. C. Gabriel, A. Aiello, W. Zhong, T. G. Euser, N. Y. Joly, P. Banzer, M. Förtsch, D. Elser, U. L. Andersen, Ch. Marquardt, P. St. J. Russell, and G. Leuchs, “Entangling different degrees of freedom by quadrature squeezing cylindrically polarized modes,” Phys. Rev. Lett. 106(6), 060502 (2011). [CrossRef] [PubMed]

]. We have exploited these findings to present a new way of visualizing CPMs on a pair of hybrid Poincaré spheres. This permits, at a glance, to easily recognize how phase retarders such as waveplates or cylindrical lenses act on these complex polarization patterns. Furthermore, we have extended our discussion of the properties of these modes to the quantum domain and have shown that the classical inseparability property leads to the intriguing feature of entanglement.

In this manuscript we have made the assumption that in the experiment, interference between the “+” and the “−” modes do not occur. However, in experimental realizations interference may be present due to birefringence and imperfect optical elements. This might result in modes with slightly asymmetric structural properties. Experimentally, care has to be taken to minimize such asymmetries, however, we believe that minor imperfections will not substantially affect the measurement outcome as it has been shown in [12

12. C. Gabriel, A. Aiello, W. Zhong, T. G. Euser, N. Y. Joly, P. Banzer, M. Förtsch, D. Elser, U. L. Andersen, Ch. Marquardt, P. St. J. Russell, and G. Leuchs, “Entangling different degrees of freedom by quadrature squeezing cylindrically polarized modes,” Phys. Rev. Lett. 106(6), 060502 (2011). [CrossRef] [PubMed]

].

In general, our results presented here give insight into the theoretical structure of these modes and their particular behaviour. We believe that the presented formalism and results are of utility to both the quantum and the classical communities especially in view of the recently growing interest in the theory and applications of light beams with complex spatial and polarization patterns.

The authors thank Peter Banzer for fruitful discussions.

8. Appendix

In our hybrid Poincaré sphere representation, only certain transformations are permitted. Due to the fact that these constraints are not intuitive, we demonstrate here how they can be derived. Generally, besides the particular cases illustrated in Fig. 8, the fundamental rules that an optical transformation must obey in order not to violate the global rotation law in Eq. (14) and consequently not to generate mixing between the co- and counter-rotating modes, can be deduced from perfectly general principles, as follows.

Consider a generic transformation 𝒯 = 𝒨 × 𝒫 that acts upon both spatial (𝒨) and polarization (𝒫) degrees of freedom of the field separately. Here, it is convenient to adopt a quantum-like notation, and write
x^|e1,y^|e2,
(85a)
ψ10(x)|ψ1(x),ψ01(x)|ψ2(x),
(85b)
where the symbol “≐” stands for: “is represented by”. Note, that the quantum-mechanical notation is just a convenient manner of writing our classical modes of the electro-magnetic field and it has nothing to do with quantum-mechanical interpretation. However, for the sake of completeness, our hybrid classical description has been extended to the quantum domain in Sec. 6. Thus, for example, our set of cylindrically polarized modes may be represented as:
uR±|uR±=12(±|ψ1|e1+|ψ2|e2),
(86)
uA±|uA±=12(|ψ2|e1+|ψ1|e2).
(87)
In passing, we note that these states have the formal structure of the quantum Bell-states. More specifically, if one establishes the following formal equivalence between the two qubits (A,B) Hilbert space and our two degrees of freedom (polarization and spatial mode) space: |0A〉 ∼ , |1A〉 ∼ ŷ, |0B〉 ∼ ψ 10, |1B〉 ∼ ψ 01, then it is not difficult to show that our basis is mathematically equivalent to the quantum Bell basis [27

27. M. A. Nielsen and I. L. Chuang, Quantum computation and quantum information (Cambridge University Press, 2000).

]:
{uR+,uA+,uR,uA}{|Φ+,|Ψ,|Φ,|Ψ+},
(88)
where |Φ±=(|0A,0B±|1A,1B)/2 and |Ψ±=(|0A,1B±|1A,0B)/2. This analogy has been recently exploited in [28

28. C. V. S. Borges, M. Hor-Meyll, J. A. O. Huguenin, and A. Z. Khoury, “Bell-like inequality for the spin-orbit separability of a laser beam,” Phys. Rev. A 82(3), 033833 (2010). [CrossRef]

].

Now, let and be such that = , the two spatial and polarization operators corresponding to the physical transformation 𝒯 = 𝒨 × 𝒫. Then, by definition such operation transforms |u ±(x)〉 into |u ±(x)〉. It is clear that this new state maintains the same symmetry under rotation of the original state |u ±(x)〉 if and only if:
G^±(φ)T^|u±(x)=T^|u±(x).
(96)
By multiplying both sides of Eq. (89) by , one straightforwardly obtains:
T^|u±(x)=T^G^±(φ)|u±(x).
(97)
Thus, by using Eq. (97) into Eq. (93) one easily obtain:
G^±(φ)T^|u±(x)=T^G^±(φ)|u±(x).
(98)
or, equivalently:
[G^±(φ)T^T^G^±(φ)]|u±(x)=0.
(99)
This equation can be easily recast in commutator form by rewriting it as follows:
[G^±(φ),T^]|u±(x)=0.
(100)
Of course, any operator that commutes with Ĝ ± (φ), namely [Ĝ ± (φ),] = 0, automatically fulfills Eq. (100). However, this condition is sufficient but not necessary because Eq. (100) represents a much weaker constraint, as it only requires that the kernel (or, at least, part of it) of the operator [Ĝ ± (φ),] coincides with the subspace spanned by either |u +(x)〉 or |u (x)〉. In other words, all what is required to by Eq. (100), is that the commutator [Ĝ ± (φ),] has a null eigenvalue in correspondence of the eigenvector |u ±(x)〉.

Now, having this caveat in mind, we can rewrite the condition (100) as follows:
[R^(φ)R^(φ)][M^P^]=[M^P^][R^(φ)R^(φ)],
(101)
or, equivalently:
R^(φ)M^[R^(φ)]1R^(φ)P^[R^(φ)]1=M^P^,
(102)
where all the operators are understood to be restricted to the subspaces spanned by |g ±(x)〉. Obviously, this commutator expression is satisfied whenever the following two conditions are separately satisfied:
[R^(φ),M^]=0,[R^(φ),P^]=0,
(103)
where the arbitrariness of the angle φ has been exploited. But, again, this condition is sufficient but not at all necessary. In fact, it is not difficult to see that if one choose either
M=(m1m2m2m1)andP=(p1p2p2p1),
(104)
or
M=(m1m2m2m1)andP=(p1p2p2p1),
(105)
with m 1, m 2, p 1, p 2 ∈ ℂ, then Eq. (100) is always satisfied but Eqs. (103) are satisfied only by the first choice (104) that represent orthogonal matrices whenever m12+m22=1=p12+p22.

It is worth to notice that a circular polarizer (left or right) is represented by the Jones matrix
P=12(1±ii1),
(106)
which is of the form (104). Conversely, a half-wave plate with fast axis at angle α with respect to the horizontal axis is represented by the Jones matrix
P=(cos(2α)sin(2α)sin(2α)cos(2α)),
(107)
which is of the form (105).

References and links

1.

A. E. Siegman, Lasers (University Science Books, 1986).

2.

R. Martinez-Herrero, P. M. Mejías, and G. Piquero, Characterization of Partially Polarized Light Fields (Springer, 2009). [CrossRef]

3.

C. J. R. Sheppard, “Polarization of almost-plane waves,” J. Opt. Soc. Am. A 17, 335 (2000). [CrossRef]

4.

C. Maurer, A. Jesacher, S. Fürhapter, S. Bernet, and M. Ritsch-Marte, “Tailoring of arbitrary optical vector beams,” N. J. Phys. 9, 78 (2007). [CrossRef]

5.

S. Quabis, R. Dorn, M. Eberler, O. Glöckl, and G. Leuchs, “Focusing light to a tighter spot,” Opt. Comm. 179, 1 – 7 (2000). [CrossRef]

6.

R. Dorn, S. Quabis, and G. Leuchs, “Sharper focus for a radially polarized light beam,” Phys. Rev. Lett. 91(23), 233901 (2003). [CrossRef] [PubMed]

7.

N. Huse, A. Schönle, and S. W. Hell, “Z-polarized confocal microscopy,” J. Biomed. Opt. 6, 273–276 (2001). [CrossRef] [PubMed]

8.

M. Sondermann, R. Maiwald, H. Konermann, N. Lindlein, U. Peschel, and G. Leuchs, “Design of a mode converter for efficient light-atom coupling in free space,” Appl. Phys. B 89, 489–492 (2007). [CrossRef]

9.

M. Meier, V. Romano, and T. Feurer, “Material processing with pulsed radially and azimuthally polarized laser radiation,” Appl. Phys. A 86, 329–334 (2006). [CrossRef]

10.

P. Banzer, U. Peschel, S. Quabis, and G. Leuchs, “On the experimental investigation of the electric and magnetic response of a single nano-structure,” Opt. Express 18(10), 10905–10923 (2010). [CrossRef] [PubMed]

11.

M. Lassen, G. Leuchs, and U. L. Andersen, “Continuous variable entanglement and squeezing of orbital angular momentum states,” Phys. Rev. Lett. 102(16), 163602 (2009). [CrossRef] [PubMed]

12.

C. Gabriel, A. Aiello, W. Zhong, T. G. Euser, N. Y. Joly, P. Banzer, M. Förtsch, D. Elser, U. L. Andersen, Ch. Marquardt, P. St. J. Russell, and G. Leuchs, “Entangling different degrees of freedom by quadrature squeezing cylindrically polarized modes,” Phys. Rev. Lett. 106(6), 060502 (2011). [CrossRef] [PubMed]

13.

J. T. Barreiro, T.-C. Wei, and P. G. Kwiat, “Remote preparation of single-photon “hybrid” entangled and vector-polarization states,” Phys. Rev. Lett. 105(3), 030407 (2010). [CrossRef] [PubMed]

14.

R. Loudon, The Quantum Theory of Light (Oxford University Press, 2000).

15.

M. J. Padgett and J. Courtial, “Poincaré-sphere equivalent for light beams containing orbital angular momentum,” Opt. Lett. 24(7), 430–432 (1999). [CrossRef]

16.

A. Peres, Quantum Theory: Concept and Methods (Kluwer Academic, Boston, 1995).

17.

M. V. Berry, “Optical currents,” J. Opt. A, Pure Appl. Opt. 11(9), 094001 (2009). [CrossRef]

18.

A. Aiello, N. Lindlein, Ch. Marquardt, and G. Leuchs, “Transverse angular momentum and geometric spin hall effect of light,” Phys. Rev. Lett. 103(10), 100401 (2009). [CrossRef] [PubMed]

19.

J. M. Jauch and F. Rohrlich, The Theory of Photons and Electrons: The relativistic Field Theory of Charged Particles with Spin one-half (Addison-Wesley Publishing Company, 1955).

20.

R. M. A. Azzam and N. Bashra, Ellipsometry and Polarized Light (North Holland Personal Library, 1988).

21.

L. Allen, M. W. Beijersbergen, R. J. C. Spreeuw, and J. P. Woerdman, “Orbital angular momentum of light and the transformation of laguerre-gaussian laser modes,” Phys. Rev. A 45(11), 8185–8189 (1992). [CrossRef] [PubMed]

22.

H. Sasada and M. Okamoto, “Transverse-mode beam splitter of a light beam and its application to quantum cryptography,” Phys. Rev. A 68(1), 012323 (2003). [CrossRef]

23.

M. T. L. Hsu, W. P. Bowen, and P. K. Lam, “Spatial-state stokes-operator squeezing and entanglement for optical beams,” Phys. Rev. A 79(4), 043825 (2009). [CrossRef]

24.

N. Korolkova, G. Leuchs, R. Loudon, T. C. Ralph, and Ch. Silberhorn, “Polarization squeezing and continuous-variable polarization entanglement,” Phys. Rev. A 65(5), 052306 (2002). [CrossRef]

25.

J. N. Damask, Polarization Optics in Telecommunication (Springer, 2005).

26.

W. P. Bowen, R. Schnabel, P. K. Lam, and T. C. Ralph, “Experimental investigation of criteria for continuous variable entanglement,” Phys. Rev. Lett. 90(4), 043601 (2003). [CrossRef] [PubMed]

27.

M. A. Nielsen and I. L. Chuang, Quantum computation and quantum information (Cambridge University Press, 2000).

28.

C. V. S. Borges, M. Hor-Meyll, J. A. O. Huguenin, and A. Z. Khoury, “Bell-like inequality for the spin-orbit separability of a laser beam,” Phys. Rev. A 82(3), 033833 (2010). [CrossRef]

OCIS Codes
(260.0260) Physical optics : Physical optics
(260.5430) Physical optics : Polarization
(270.0270) Quantum optics : Quantum optics

ToC Category:
Physical Optics

History
Original Manuscript: December 20, 2010
Revised Manuscript: February 18, 2011
Manuscript Accepted: February 22, 2011
Published: May 4, 2011

Citation
Annemarie Holleczek, Andrea Aiello, Christian Gabriel, Christoph Marquardt, and Gerd Leuchs, "Classical and quantum properties of cylindrically polarized states of light," Opt. Express 19, 9714-9736 (2011)
http://www.opticsinfobase.org/oe/abstract.cfm?URI=oe-19-10-9714


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References

  1. A. E. Siegman, Lasers (University Science Books, 1986).
  2. R. Martinez-Herrero, P. M. Mejías, and G. Piquero, Characterization of Partially Polarized Light Fields (Springer, 2009). [CrossRef]
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