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Zeeman effect



The Zeeman effect (/ˈzmən/; Dutch pronunciation: [ˈzeːmɑn]), named after Dutch physicist Pieter Zeeman, is the effect of splitting of a spectral line into several components in the presence of a static magnetic field. It is analogous to the Stark effect, the splitting of a spectral line into several components in the presence of an electric field. Also similar to the Stark effect, transitions between different components have, in general, different intensities, with some being entirely forbidden (in the dipole approximation), as governed by the selection rules.

The spectral lines of mercury vapor lamp at wavelength 546.1 nm, showing anomalous Zeeman effect. (A) Without magnetic field. (B) With magnetic field, spectral lines split as transverse Zeeman effect. (C) With magnetic field, split as longitudinal Zeeman effect. The spectral lines were obtained using a Fabry–Pérot interferometer.
Zeeman splitting of the 5s level of 87Rb, including fine structure and hyperfine structure splitting. Here F = J + I, where I is the nuclear spin (for 87Rb, I = 32).

Since the distance between the Zeeman sub-levels is a function of magnetic field strength, this effect can be used to measure magnetic field strength, e.g. that of the Sun and other stars or in laboratory plasmas. The Zeeman effect is very important in applications such as nuclear magnetic resonance spectroscopy, electron spin resonance spectroscopy, magnetic resonance imaging (MRI) and Mössbauer spectroscopy. It may also be utilized to improve accuracy in atomic absorption spectroscopy. A theory about the magnetic sense of birds assumes that a protein in the retina is changed due to the Zeeman effect.[1]

When the spectral lines are absorption lines, the effect is called inverse Zeeman effect.

Contents


Nomenclature

Historically, one distinguishes between the normal and an anomalous Zeeman effect (discovered by Thomas Preston in Dublin, Ireland[2]). The anomalous effect appears on transitions where the net spin of the electrons is non-zero. It was called "anomalous" because the electron spin had not yet been discovered, and so there was no good explanation for it at the time that Zeeman observed the effect.

At higher magnetic field strength the effect ceases to be linear. At even higher field strengths, comparable to the strength of the atom's internal field, the electron coupling is disturbed and the spectral lines rearrange. This is called the Paschen–Back effect.

In the modern scientific literature, these terms are rarely used, with a tendency to use just the "Zeeman effect".


Theoretical presentation

The total Hamiltonian of an atom in a magnetic field is

\({\displaystyle H=H_{0}+V_{\rm {M}},\ }\)

where \({\displaystyle H_{0}}\) is the unperturbed Hamiltonian of the atom, and \({\displaystyle V_{\rm {M}}}\) is the perturbation due to the magnetic field:

\({\displaystyle V_{\rm {M}}=-{\vec {\mu }}\cdot {\vec {B}},}\)

where \({\displaystyle {\vec {\mu }}}\) is the magnetic moment of the atom. The magnetic moment consists of the electronic and nuclear parts; however, the latter is many orders of magnitude smaller and will be neglected here. Therefore,

\({\displaystyle {\vec {\mu }}\approx -{\frac {\mu _{\rm {B}}g{\vec {J}}}{\hbar }},}\)

where \({\displaystyle \mu _{\rm {B}}}\) is the Bohr magneton, \({\displaystyle {\vec {J}}}\) is the total electronic angular momentum, and \({\displaystyle g}\) is the Landé g-factor. A more accurate approach is to take into account that the operator of the magnetic moment of an electron is a sum of the contributions of the orbital angular momentum \({\displaystyle {\vec {L}}}\) and the spin angular momentum \({\displaystyle {\vec {S}}}\), with each multiplied by the appropriate gyromagnetic ratio:

\({\displaystyle {\vec {\mu }}=-{\frac {\mu _{\rm {B}}(g_{l}{\vec {L}}+g_{s}{\vec {S}})}{\hbar }},}\)

where \({\displaystyle g_{l}=1}\) and \({\displaystyle g_{s}\approx 2.0023192}\) (the latter is called the anomalous gyromagnetic ratio; the deviation of the value from 2 is due to the effects of quantum electrodynamics). In the case of the LS coupling, one can sum over all electrons in the atom:

\({\displaystyle g{\vec {J}}=\left\langle \sum _{i}(g_{l}{\vec {l_{i}}}+g_{s}{\vec {s_{i}}})\right\rangle =\left\langle (g_{l}{\vec {L}}+g_{s}{\vec {S}})\right\rangle ,}\)

where \({\displaystyle {\vec {L}}}\) and \({\displaystyle {\vec {S}}}\) are the total orbital momentum and spin of the atom, and averaging is done over a state with a given value of the total angular momentum.

If the interaction term \({\displaystyle V_{M}}\) is small (less than the fine structure), it can be treated as a perturbation; this is the Zeeman effect proper. In the Paschen–Back effect, described below, \({\displaystyle V_{M}}\) exceeds the LS coupling significantly (but is still small compared to \({\displaystyle H_{0}}\)). In ultra-strong magnetic fields, the magnetic-field interaction may exceed \({\displaystyle H_{0}}\), in which case the atom can no longer exist in its normal meaning, and one talks about Landau levels instead. There are intermediate cases which are more complex than these limit cases.


Weak field (Zeeman effect)

If the spin–orbit interaction dominates over the effect of the external magnetic field, \({\displaystyle \scriptstyle {\vec {L}}}\) and \({\displaystyle \scriptstyle {\vec {S}}}\) are not separately conserved, only the total angular momentum \({\displaystyle \scriptstyle {\vec {J}}={\vec {L}}+{\vec {S}}}\) is. The spin and orbital angular momentum vectors can be thought of as precessing about the (fixed) total angular momentum vector \({\displaystyle \scriptstyle {\vec {J}}}\). The (time-)"averaged" spin vector is then the projection of the spin onto the direction of \({\displaystyle \scriptstyle {\vec {J}}}\):

\({\displaystyle {\vec {S}}_{\rm {avg}}={\frac {({\vec {S}}\cdot {\vec {J}})}{J^{2}}}{\vec {J}}}\)

and for the (time-)"averaged" orbital vector:

\({\displaystyle {\vec {L}}_{\rm {avg}}={\frac {({\vec {L}}\cdot {\vec {J}})}{J^{2}}}{\vec {J}}.}\)

Thus,

\({\displaystyle \langle V_{\rm {M}}\rangle ={\frac {\mu _{\rm {B}}}{\hbar }}{\vec {J}}\left(g_{L}{\frac {{\vec {L}}\cdot {\vec {J}}}{J^{2}}}+g_{S}{\frac {{\vec {S}}\cdot {\vec {J}}}{J^{2}}}\right)\cdot {\vec {B}}.}\)

Using \({\displaystyle \scriptstyle {\vec {L}}={\vec {J}}-{\vec {S}}}\) and squaring both sides, we get

\({\displaystyle {\vec {S}}\cdot {\vec {J}}={\frac {1}{2}}(J^{2}+S^{2}-L^{2})={\frac {\hbar ^{2}}{2}}[j(j+1)-l(l+1)+s(s+1)],}\)

and: using \({\displaystyle \scriptstyle {\vec {S}}={\vec {J}}-{\vec {L}}}\) and squaring both sides, we get

\({\displaystyle {\vec {L}}\cdot {\vec {J}}={\frac {1}{2}}(J^{2}-S^{2}+L^{2})={\frac {\hbar ^{2}}{2}}[j(j+1)+l(l+1)-s(s+1)].}\)

Combining everything and taking \({\displaystyle \scriptstyle J_{z}=\hbar m_{j}}\), we obtain the magnetic potential energy of the atom in the applied external magnetic field,

\({\displaystyle {\begin{aligned}V_{\rm {M}}&=\mu _{\rm {B}}Bm_{j}\left[g_{L}{\frac {j(j+1)+l(l+1)-s(s+1)}{2j(j+1)}}+g_{S}{\frac {j(j+1)-l(l+1)+s(s+1)}{2j(j+1)}}\right]\\&=\mu _{\rm {B}}Bm_{j}\left[1+(g_{S}-1){\frac {j(j+1)-l(l+1)+s(s+1)}{2j(j+1)}}\right],\\&=\mu _{\rm {B}}Bm_{j}g_{j}\end{aligned}}}\)

where the quantity in square brackets is the Landé g-factor gJ of the atom (\({\displaystyle g_{L}=1}\) and \({\displaystyle g_{S}\approx 2}\)) and \({\displaystyle m_{j}}\) is the z-component of the total angular momentum. For a single electron above filled shells \({\displaystyle s=1/2}\) and \({\displaystyle j=l\pm s}\), the Landé g-factor can be simplified into:

\({\displaystyle g_{j}=1\pm {\frac {g_{S}-1}{2l+1}}}\)

Taking \({\displaystyle V_{m}}\) to be the perturbation, the Zeeman correction to the energy is

\({\displaystyle {\begin{aligned}E_{\rm {Z}}^{(1)}=\langle nljm_{j}|H_{\rm {Z}}^{'}|nljm_{j}\rangle =\langle V_{M}\rangle _{\Psi }=\mu _{\rm {B}}g_{J}B_{\rm {ext}}m_{j}\end{aligned}}}\)

Example: Lyman-alpha transition in hydrogen

The Lyman-alpha transition in hydrogen in the presence of the spin–orbit interaction involves the transitions

\({\displaystyle 2P_{1/2}\to 1S_{1/2}}\) and \({\displaystyle 2P_{3/2}\to 1S_{1/2}.}\)

In the presence of an external magnetic field, the weak-field Zeeman effect splits the 1S1/2 and 2P1/2 levels into 2 states each (\({\displaystyle m_{j}=1/2,-1/2}\)) and the 2P3/2 level into 4 states (\({\displaystyle m_{j}=3/2,1/2,-1/2,-3/2}\)). The Landé g-factors for the three levels are:

\({\displaystyle g_{J}=2}\) for \({\displaystyle 1S_{1/2}}\) (j=1/2, l=0)
\({\displaystyle g_{J}=2/3}\) for \({\displaystyle 2P_{1/2}}\) (j=1/2, l=1)
\({\displaystyle g_{J}=4/3}\) for \({\displaystyle 2P_{3/2}}\) (j=3/2, l=1).

Note in particular that the size of the energy splitting is different for the different orbitals, because the gJ values are different. On the left, fine structure splitting is depicted. This splitting occurs even in the absence of a magnetic field, as it is due to spin–orbit coupling. Depicted on the right is the additional Zeeman splitting, which occurs in the presence of magnetic fields.

Possible transitions for the weak Zeeman effect
Initial state

(\({\displaystyle n=2,l=1}\))

\({\displaystyle \mid j,m_{j}\rangle }\)

Final state

(\({\displaystyle n=1,l=0}\))

\({\displaystyle \mid j,m_{j}\rangle }\)

Energy perturbation
\({\displaystyle \left|{\frac {1}{2}},\pm {\frac {1}{2}}\right\rangle }\) \({\displaystyle \left|{\frac {1}{2}},\pm {\frac {1}{2}}\right\rangle }\) \({\displaystyle \mp {\frac {2}{3}}\mu _{\rm {B}}B}\)
\({\displaystyle \left|{\frac {1}{2}},\pm {\frac {1}{2}}\right\rangle }\) \({\displaystyle \left|{\frac {1}{2}},\mp {\frac {1}{2}}\right\rangle }\) \({\displaystyle \pm {\frac {4}{3}}\mu _{\rm {B}}B}\)
\({\displaystyle \left|{\frac {3}{2}},\pm {\frac {3}{2}}\right\rangle }\) \({\displaystyle \left|{\frac {1}{2}},\pm {\frac {1}{2}}\right\rangle }\) \({\displaystyle \pm \mu _{\rm {B}}B}\)
\({\displaystyle \left|{\frac {3}{2}},\pm {\frac {1}{2}}\right\rangle }\) \({\displaystyle \left|{\frac {1}{2}},\pm {\frac {1}{2}}\right\rangle }\) \({\displaystyle \mp {\frac {1}{3}}\mu _{\rm {B}}B}\)
\({\displaystyle \left|{\frac {3}{2}},\pm {\frac {1}{2}}\right\rangle }\) \({\displaystyle \left|{\frac {1}{2}},\mp {\frac {1}{2}}\right\rangle }\) \({\displaystyle \pm {\frac {5}{3}}\mu _{\rm {B}}B}\)

Strong field (Paschen–Back effect)

The Paschen–Back effect is the splitting of atomic energy levels in the presence of a strong magnetic field. This occurs when an external magnetic field is sufficiently strong to disrupt the coupling between orbital (\({\displaystyle {\vec {L}}}\)) and spin (\({\displaystyle {\vec {S}}}\)) angular momenta. This effect is the strong-field limit of the Zeeman effect. When \({\displaystyle s=0}\), the two effects are equivalent. The effect was named after the German physicists Friedrich Paschen and Ernst E. A. Back.[3]

When the magnetic-field perturbation significantly exceeds the spin–orbit interaction, one can safely assume \({\displaystyle [H_{0},S]=0}\). This allows the expectation values of \({\displaystyle L_{z}}\) and \({\displaystyle S_{z}}\) to be easily evaluated for a state \({\displaystyle |\psi \rangle }\). The energies are simply

\({\displaystyle E_{z}=\left\langle \psi \left|H_{0}+{\frac {B_{z}\mu _{\rm {B}}}{\hbar }}(L_{z}+g_{s}S_{z})\right|\psi \right\rangle =E_{0}+B_{z}\mu _{\rm {B}}(m_{l}+g_{s}m_{s}).}\)

The above may be read as implying that the LS-coupling is completely broken by the external field. However \({\displaystyle m_{l}}\) and \({\displaystyle m_{s}}\) are still "good" quantum numbers. Together with the selection rules for an electric dipole transition, i.e., \({\displaystyle \Delta s=0,\Delta m_{s}=0,\Delta l=\pm 1,\Delta m_{l}=0,\pm 1}\) this allows to ignore the spin degree of freedom altogether. As a result, only three spectral lines will be visible, corresponding to the \({\displaystyle \Delta m_{l}=0,\pm 1}\) selection rule. The splitting \({\displaystyle \Delta E=B\mu _{\rm {B}}\Delta m_{l}}\) is independent of the unperturbed energies and electronic configurations of the levels being considered. In general (if \({\displaystyle s\neq 0}\)), these three components are actually groups of several transitions each, due to the residual spin–orbit coupling.

In general, one must now add spin–orbit coupling and relativistic corrections (which are of the same order, known as 'fine structure') as a perturbation to these 'unperturbed' levels. First order perturbation theory with these fine-structure corrections yields the following formula for the hydrogen atom in the Paschen–Back limit:[4]

\({\displaystyle E_{z+fs}=E_{z}+{\frac {m_{e}c^{2}\alpha ^{4}}{2n^{3}}}\left\{{\frac {3}{4n}}-\left[{\frac {l(l+1)-m_{l}m_{s}}{l(l+1/2)(l+1)}}\right]\right\}.}\)
Possible Lyman-alpha transitions for the strong regime
Initial state

(\({\displaystyle n=2,l=1}\))

\({\displaystyle \mid m_{l},m_{s}\rangle }\)

Initial energy Perturbation Final state

(\({\displaystyle n=1,l=0}\))

\({\displaystyle \mid m_{l},m_{s}\rangle }\)

\({\displaystyle \left|1,{\frac {1}{2}}\right\rangle }\) \({\displaystyle \pm 2\mu _{\rm {B}}B_{z}}\) \({\displaystyle \left|0,{\frac {1}{2}}\right\rangle }\)
\({\displaystyle \left|0,{\frac {1}{2}}\right\rangle }\) \({\displaystyle +\mu _{\rm {B}}B_{z}}\) \({\displaystyle \left|0,{\frac {1}{2}}\right\rangle }\)
\({\displaystyle \left|1,-{\frac {1}{2}}\right\rangle }\) \({\displaystyle 0}\) \({\displaystyle \left|0,-{\frac {1}{2}}\right\rangle }\)
\({\displaystyle \left|-1,{\frac {1}{2}}\right\rangle }\) \({\displaystyle 0}\) \({\displaystyle \left|0,{\frac {1}{2}}\right\rangle }\)
\({\displaystyle \left|0,-{\frac {1}{2}}\right\rangle }\) \({\displaystyle -\mu _{\rm {B}}B_{z}}\) \({\displaystyle \left|0,-{\frac {1}{2}}\right\rangle }\)
\({\displaystyle \left|-1,-{\frac {1}{2}}\right\rangle }\) \({\displaystyle -2\mu _{\rm {B}}B_{z}}\) \({\displaystyle \left|0,-{\frac {1}{2}}\right\rangle }\)

Intermediate field for j = 1/2

In the magnetic dipole approximation, the Hamiltonian which includes both the hyperfine and Zeeman interactions is

\({\displaystyle H=hA{\vec {I}}\cdot {\vec {J}}-{\vec {\mu }}\cdot {\vec {B}}}\)
\({\displaystyle H=hA{\vec {I}}\cdot {\vec {J}}+(\mu _{\rm {B}}g_{J}{\vec {J}}+\mu _{\rm {N}}g_{I}{\vec {I}})\cdot {\vec {\rm {B}}}}\)

where \({\displaystyle A}\) is the hyperfine splitting (in Hz) at zero applied magnetic field, \({\displaystyle \mu _{\rm {B}}}\) and \({\displaystyle \mu _{\rm {N}}}\) are the Bohr magneton and nuclear magneton respectively, \({\displaystyle {\vec {J}}}\) and \({\displaystyle {\vec {I}}}\) are the electron and nuclear angular momentum operators and \({\displaystyle g_{J}}\) is the Landé g-factor:

\({\displaystyle g_{J}=g_{L}{\frac {J(J+1)+L(L+1)-S(S+1)}{2J(J+1)}}+g_{S}{\frac {J(J+1)-L(L+1)+S(S+1)}{2J(J+1)}}}\).

In the case of weak magnetic fields, the Zeeman interaction can be treated as a perturbation to the \({\displaystyle |F,m_{f}\rangle }\) basis. In the high field regime, the magnetic field becomes so strong that the Zeeman effect will dominate, and one must use a more complete basis of \({\displaystyle |I,J,m_{I},m_{J}\rangle }\) or just \({\displaystyle |m_{I},m_{J}\rangle }\) since \({\displaystyle I}\) and \({\displaystyle J}\) will be constant within a given level.

To get the complete picture, including intermediate field strengths, we must consider eigenstates which are superpositions of the \({\displaystyle |F,m_{F}\rangle }\) and \({\displaystyle |m_{I},m_{J}\rangle }\) basis states. For \({\displaystyle J=1/2}\), the Hamiltonian can be solved analytically, resulting in the Breit–Rabi formula. Notably, the electric quadrupole interaction is zero for \({\displaystyle L=0}\) (\({\displaystyle J=1/2}\)), so this formula is fairly accurate.

We now utilize quantum mechanical ladder operators, which are defined for a general angular momentum operator \({\displaystyle L}\) as

\({\displaystyle L_{\pm }\equiv L_{x}\pm iL_{y}}\)

These ladder operators have the property

\({\displaystyle L_{\pm }|L_{,}m_{L}\rangle ={\sqrt {(L\mp m_{L})(L\pm m_{L}+1)}}|L,m_{L}\pm 1\rangle }\)

as long as \({\displaystyle m_{L}}\) lies in the range \({\displaystyle {-L,\dots ...,L}}\) (otherwise, they return zero). Using ladder operators \({\displaystyle J_{\pm }}\) and \({\displaystyle I_{\pm }}\) We can rewrite the Hamiltonian as

\({\displaystyle H=hAI_{z}J_{z}+{\frac {hA}{2}}(J_{+}I_{-}+J_{-}I_{+})+\mu _{\rm {B}}Bg_{J}J_{z}+\mu _{\rm {N}}Bg_{I}I_{z}}\)

We can now see that at all times, the total angular momentum projection \({\displaystyle m_{F}=m_{J}+m_{I}}\) will be conserved. This is because both \({\displaystyle J_{z}}\) and \({\displaystyle I_{z}}\) leave states with definite \({\displaystyle m_{J}}\) and \({\displaystyle m_{I}}\) unchanged, while \({\displaystyle J_{+}I_{-}}\) and \({\displaystyle J_{-}I_{+}}\) either increase \({\displaystyle m_{J}}\) and decrease \({\displaystyle m_{I}}\) or vice versa, so the sum is always unaffected. Furthermore, since \({\displaystyle J=1/2}\) there are only two possible values of \({\displaystyle m_{J}}\) which are \({\displaystyle \pm 1/2}\). Therefore, for every value of \({\displaystyle m_{F}}\) there are only two possible states, and we can define them as the basis:

\({\displaystyle |\pm \rangle \equiv |m_{J}=\pm 1/2,m_{I}=m_{F}\mp 1/2\rangle }\)

This pair of states is a Two-level quantum mechanical system. Now we can determine the matrix elements of the Hamiltonian:

\({\displaystyle \langle \pm |H|\pm \rangle =-{\frac {1}{4}}hA+\mu _{\rm {N}}Bg_{I}m_{F}\pm {\frac {1}{2}}(hAm_{F}+\mu _{\rm {B}}Bg_{J}-\mu _{\rm {N}}Bg_{I}))}\)
\({\displaystyle \langle \pm |H|\mp \rangle ={\frac {1}{2}}hA{\sqrt {(I+1/2)^{2}-m_{F}^{2}}}}\)

Solving for the eigenvalues of this matrix, (as can be done by hand - see Two-level quantum mechanical system, or more easily, with a computer algebra system) we arrive at the energy shifts:

\({\displaystyle \Delta E_{F=I\pm 1/2}=-{\frac {h\Delta W}{2(2I+1)}}+\mu _{\rm {N}}g_{I}m_{F}B\pm {\frac {h\Delta W}{2}}{\sqrt {1+{\frac {2m_{F}x}{I+1/2}}+x^{2}}}}\)
\({\displaystyle x\equiv {\frac {B(\mu _{\rm {B}}g_{J}-\mu _{\rm {N}}g_{I})}{h\Delta W}}\quad \quad \Delta W=A\left(I+{\frac {1}{2}}\right)}\)

where \({\displaystyle \Delta W}\) is the splitting (in units of Hz) between two hyperfine sublevels in the absence of magnetic field \({\displaystyle B}\), \({\displaystyle x}\) is referred to as the 'field strength parameter' (Note: for \({\displaystyle m_{F}=\pm (I+1/2)}\) the expression under the square root is an exact square, and so the last term should be replaced by \({\displaystyle +{\frac {h\Delta W}{2}}(1\pm x)}\)). This equation is known as the Breit–Rabi formula and is useful for systems with one valence electron in an \({\displaystyle s}\) (\({\displaystyle J=1/2}\)) level.[5][6]

Note that index \({\displaystyle F}\) in \({\displaystyle \Delta E_{F=I\pm 1/2}}\) should be considered not as total angular momentum of the atom but as asymptotic total angular momentum. It is equal to total angular momentum only if \({\displaystyle B=0}\) otherwise eigenvectors corresponding different eigenvalues of the Hamiltonian are the superpositions of states with different \({\displaystyle F}\) but equal \({\displaystyle m_{F}}\) (the only exceptions are \({\displaystyle |F=I+1/2,m_{F}=\pm F\rangle }\)).


Applications

Astrophysics

Zeeman effect on a sunspot spectral line

George Ellery Hale was the first to notice the Zeeman effect in the solar spectra, indicating the existence of strong magnetic fields in sunspots. Such fields can be quite high, on the order of 0.1 tesla or higher. Today, the Zeeman effect is used to produce magnetograms showing the variation of magnetic field on the sun.

Laser cooling

The Zeeman effect is utilized in many laser cooling applications such as a magneto-optical trap and the Zeeman slower.

Zeeman-energy mediated coupling of spin and orbital motions

Spin–orbit interaction in crystals is usually attributed to coupling of Pauli matrices \({\displaystyle {\vec {\sigma }}}\) to electron momentum \({\displaystyle {\vec {k}}}\) which exists even in the absence of magnetic field \({\displaystyle {\vec {B}}}\). However, under the conditions of the Zeeman effect, when \({\displaystyle {\vec {B}}\neq 0}\), a similar interaction can be achieved by coupling \({\displaystyle {\vec {\sigma }}}\) to the electron coordinate \({\displaystyle {\vec {r}}}\) through the spatially inhomogeneous Zeeman Hamiltonian

\({\displaystyle H_{\rm {Z}}={\frac {1}{2}}({\vec {B}}{\hat {g}}{\vec {\sigma }})}\),

where \({\displaystyle {\hat {g}}}\) is a tensorial Landé g-factor and either \({\displaystyle {\vec {B}}={\vec {B}}({\vec {r}})}\) or \({\displaystyle {\hat {g}}={\hat {g}}({\vec {r}})}\), or both of them, depend on the electron coordinate \({\displaystyle {\vec {r}}}\). Such \({\displaystyle {\vec {r}}}\)-dependent Zeeman Hamiltonian \({\displaystyle H_{\rm {Z}}({\vec {r}})}\) couples electron spin \({\displaystyle {\vec {\sigma }}}\) to the operator \({\displaystyle {\vec {r}}}\) representing electron's orbital motion. Inhomogeneous field \({\displaystyle {\vec {B}}({\vec {r}})}\) may be either a smooth field of external sources or fast-oscillating microscopic magnetic field in antiferromagnets.[7] Spin–orbit coupling through macroscopically inhomogeneous field \({\displaystyle {\vec {B}}({\vec {r}})}\) of nanomagnets is used for electrical operation of electron spins in quantum dots through electric dipole spin resonance,[8] and driving spins by electric field due to inhomogeneous \({\displaystyle {\hat {g}}({\vec {r}})}\) has been also demonstrated.[9]


See also


References

  1. ^ Thalau, Peter; Ritz, Thorsten; Burda, Hynek; Wegner, Regina E.; Wiltschko, Roswitha (18 April 2006). "The magnetic compass mechanisms of birds and rodents are based on different physical principles" . Journal of the Royal Society Interface. 3 (9): 583–587. doi:10.1098/rsif.2006.0130 . PMC 1664646 . PMID 16849254 .
  2. ^ Preston, Thomas (1898). "Radiation phenomena in a strong magnetic field" . The Scientific Transactions of the Royal Dublin Society. 2nd series. 6: 385–342.
  3. ^ Paschen, F.; Back, E. (1921). "Liniengruppen magnetisch vervollständigt" [Line groups magnetically completed [i.e., completely resolved]]. Physica (in German). 1: 261–273. Available at: Leiden University (Netherlands)
  4. ^ Griffiths, David J. (2004). Introduction to Quantum Mechanics (2nd ed.). Prentice Hall. p. 247. ISBN 0-13-111892-7. OCLC 40251748 .
  5. ^ Woodgate, Gordon Kemble (1980). Elementary Atomic Structure (2nd ed.). Oxford, England: Oxford University Press. pp. 193–194.
  6. ^ First appeared in: Breit, G.; Rabi, I.I. (1931). "Measurement of nuclear spin". Physical Review. 38 (11): 2082–2083. Bibcode:1931PhRv...38.2082B . doi:10.1103/PhysRev.38.2082.2 .
  7. ^ S. I. Pekar and E. I. Rashba, Combined resonance in crystals in inhomogeneous magnetic fields, Sov. Phys. - JETP 20, 1295 (1965) http://www.jetp.ac.ru/cgi-bin/dn/e_020_05_1295.pdf
  8. ^ Y. Tokura, W. G. van der Wiel, T. Obata, and S. Tarucha, Coherent single electron spin control in a slanting Zeeman field, Phys. Rev. Lett. 96, 047202 (2006)
  9. ^ Salis G, Kato Y, Ensslin K, Driscoll DC, Gossard AC, Awschalom DD (2001). "Electrical control of spin coherence in semiconductor nanostructures" . Nature. 414 (6864): 619–622. doi:10.1038/414619a . PMID 11740554 . S2CID 4393582 .CS1 maint: uses authors parameter (link)

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