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NEW MAGNETIC MATERIALS BASED ON SINGLE MOLECULE MAGNETS

Sergey M. Aldoshin†, Andrew V. Palii†,‡, Denis V. Korchagin†, Boris S. Tsukerblat*

† Institute of Problems of Chemical Physics, Chernogolovka, Moscow Region, Russia ‡ Institute of Applied Physics, Academy of Sciences of Moldova, Chisinau, Moldova *Department of Chemistry, Ben-Gurion University of the Negev, Beer-Sheva, Israel

ABSTRACT

Here we briefly review some new magnetic materials based single molecule magnets con-taining transition and rare-earth ions. Within this broad theme the emphasis is made on the molecules which exhibit strong magnetic anisotropy originating from the unquenched orbital angular momentum. Along with the general concepts we consider selected examples of the systems comprising orbitally-degenerate metal ions and demonstrate how one can benefit from strong single-ion anisotropy. The role of crystal fields, spin-orbit coupling and structural factors is discussed. Some observation stemming from the in-depth analysis of the exchange interactions are summarized as a set of guiding rules for the controlled design of single mole-cule magnets exhibiting high magnetic barrier and blocking temperature.

INTRODUCTION Last years single-molecule magnets (SMMs) have been in focus of the molecular mag-

netism because of their rich physical properties and fascinating potential applications [1]. Thus, these molecules exhibit a superparamagnetic blocking characterized by slow relaxation of magnetization at temperatures lower than blocking temperature. As distinguished from the bulk magnets, they show also the magnetic hysteresis which represents an essentially single-molecule phenomenon promising for the high-density data storage [2]. On the other hand, SMMs show quantum phenomena, like quantum tunnelling of the magnetization [1,3], and also quantum coherence and interference which open up a route for using SMMs in spintronic devices and quantum computing [4,5,6,7,8,9].

The basis of the SMM area was formed by the discovery of the magnetic bistability in the so-called Mn12 cluster [1] which became the main object of the research for a long time. The first generation of SMMs was based on polynuclear magnetic complexes of 3d-ions in which strong isotropic exchange coupling between transition-metal ions leads to a high-spin ground state, S, well separated from the excited multiplets, and a negative uniaxial magnetic anisotropy, described by the conventionally accepted zero-field splitting (ZFS) spin Hamilto-nian 2ˆ

ZFS S ZH D S= . Providing DS < 0 (negative magnetic anisotropy) the ground state with the spin S is split into sublevels labelled by the full spin projections ±MS that can be represented as a barrier for spin reversal [1] as shown in Fig.1 for Mn12 cluster, S = 10. The presence of the barrier implies that the magnetic anisotropy defined as χ⊥ ‒ χ|| is negative and therefore providing DS < 0, the spin quantization axis Z is an easy axis of magnetization. It should be emphasized that the metal ions incorporated in such type of SMMs are assumed to be orbitally non-degenerate and their spins are coupled to give a certain full spin S. Consequently, the first-order orbital angular momenta are quenched in the constituent ions as well in the system entire. These non-degenerate systems may be conventionally termed “spin-type SMMs”.

Although strong efforts have been made towards practical realization of the memory units, presently the relaxation time is not long enough to store information in the required timescale. That is why the central problem in the design of SMMs with

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Fig. 1. Energy as function of the magnetization (barrier) Z B Sg Mµ µ= (g= 2).

Fig. 2. Spin dependence of the ZFS parameter DS in clusters .

high blocking temperatures is to increase the barrier for the reversal of magnetization. Alt-hough ∆b = DS S2 implies that this can be achieved by the increase of the nuclearity which would have large values of S in the ground state, A more detailed analysis shows that the ZFS parameter DS proves to be proportional to S−2 [10,11,12]. This is seen from Fig. 2 in which the parameter |DS| is shown as function of S for a series of high-spin polynuclear 3d metal clusters Mn6 [13], Mn4 [14], Mn12 [15], Mn18 [16], Mn25 [17], Mn19 [18], and Fe42[19].

As a result the anisotropy barrier ∆b does grow as S2 with the increase of the full spin S [10,11,12]. In fact, the first and most famous SMM, the so-called Mn12, exhibiting effective barrier of about 45 cm-1 and hysteresis up to 4K is still among the best spin-type SMMs in spite of the fact that much bigger spin-type clusters with higher values of the ground state spin have been synthesized (see Refs. [16,17,18,19,20,21,22]).

A promising approach is based on the idea to incorporate metal ions with first order or-bital angular momentum. This makes it possible to benefit from the first-order single-ion ani-sotropy which is definitely stronger than that in spin-clusters. A second important advantage of such systems arises from the exchange anisotropy of orbitally degenerate ions. This kind of the so-called orbitally-dependent superexchange may considerably increase the barrier. The aim of this article is to review some recent theoretical ideas underlying the phenomenon of single molecule magnetism in systems based on 3d, 4d, 5d and 4f ions and to provide some simple guiding concepts which could be of help for the rational design of SMMs of this type.

1. SINGLE-MOLECULE MAGNETS COMPRISING TRANSITION METAL IONS WITH UNQUENCHED ORBITAL ANGULAR MOMENTA

In this Section will consider examples of SMMs based on d-metal ions with unquenched orbital angular momentum, mainly the cyanide based clusters [22-37]. We will demonstrate that the conventional spin-Hamiltonian is insufficient for the description of the magnetic be-haviour of such systems and discuss more sophisticated models.

1.1 Magnetization reversal barrier in clusters containing orbitally-degenerate 3d-ions: first-order single ion anisotropy

One of the first well-documented example of such SMM was the trigonal bipyramidal cyano-bridged cluster [MnIII(CN)6]2[MnII(tmphen)2]3 (tmphen = 3, 4, 7, 8–tetramethyl–1,10–phenanthroline) [23]. The molecular geometry of this Mn5-cyanide cluster is shown in Fig. 3a. Five Mn ions form a trigonal bipyramid in which two Mn(III) ions (1 and 2) occupy the apical positions and possess the low-spin ground terms 3T1g(t2g

4) in a strong crystal field (CF) of the carbon octahedral surroundings (Fig. 3b), while the three Mn(II) ions (3, 4, 5) reside in the equatorial plane and have the high-spin ground terms ( )6 3 2

1 2g g gA t e in a weak CFs induced by the nitrogen octahedra (Fig. 3c). In addition to the spin value S = 1 the ground term 3T1g(t2g

4)

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of the Mn(III) ion carries first order orbital angular momentum. In fact the orbital triplet 3T1g is in some sense equivalent to the atomic triply degenerate P state which can be associated with L = 1. One can state that the Mn5-cyanide system drastically differs from the pure spin SMMs in which all constituent metal ions are characterized only by spin values. The presence of the unquenched orbital angular momentum can lead under some conditions to the appear-ance of rather strong uniaxial magnetic anisotropy which is responsible for the formation of the barrier for the reversal of magnetization [24,25]. It is important to note that the presence of the unquenched orbital angular momentum in the ground term created by perfect octahedral ligand field does not carry magnetic anisotropy. Fortunately, the carbon surroundings of the Mn(III) ions are distorted along the trigonal Z axis (Fig. 3a). Instead of considering the real situation occurring in the Mn5-cyanide cluster for which this trigonal distortion is rather weak, we will discuss here a hypothetical situation of rather strong trigonal distortion of the nearest surrounding of the Mn(III) ion. The analysis of such strong trigonal CF is quite instructive because it reveals with utmost clarity the role of orbital angular momentum in the formation of the barrier for magnetization reversal.

Fig. 3. Molecular structure of the Mn5-cyanide cluster (a) and electronic configurations and ground terms of the Mn(III) (b) and Mn(II) (c) ions.

The trigonal CF acting on the Mn(III) ion (which is actually axial) can be described by the Hamiltonian:

( )2 1ˆ ˆΔ 1 ,3ax ZH L L L = − +

(1)

where Δ is the axial CF parameter, ˆZL is the Z-component of the orbital angular momentum

operator and for the orbital triplet one should adopt L=1. The trigonal field splits the 3T1g state into the orbital singlet 3

2gA ( 0LM = ) and the orbital doublet 3gE ( 1LM = ± ) which is the ground

state providing 0.∆ < We will focus on the case of 3

gE ground term, which carries a residual orbital angular momentum. The splitting of the cu-bic 3T1g – term (L = 1, S = 1) is shown in Fig. 4b.

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Fig. 4. Splitting of the ground cubic 3T1g - term (a) of the Mn(III) ion by the negative trigonal CF (b) and SOC (c) in the limit of strong trigonal CF.

In case of perfect octahedral surrounding of this ion the spin-orbit coupling (SOC) acts within the cubic 3T1g (S = 1, L = 1) term and is described by the following Hamiltonian:

( ) ˆˆ ˆ1, 1SOCH S L κ λ= = = − L S ,

(2) whereλ is the SOC parameter that is negative in more than half-filled t2g – shell (Fig. 3b). The parameter κ is the orbital reduction factor (κ ≤1) accounting for the effects of covalence. The sign “minus” in the SOC Hamiltonian and also in the orbital part, of the Zeeman interac-tion appears due to the so-called T-P analogy [38]. The SOC Hamiltonian acquires the follow-ing axial form:

( ) ˆˆ ˆ1, 1 | | .SOC L Z ZH S M L Sκ λ= = ± =

(3) This set of levels (Fig. 4c) includes the ground non-Kramers doublet with

1, 1L SM M= ± = and two non-Kramers doublets with 1, 0L SM M= ± = and 1, 1L SM M= ± = ± . A remarkable feature of the ground non-Kramers doublet of Mn(III) ion is that this state is magnetic in spite of the fact that the total angular momentum projection for this state is van-ishing ( 0L SM M+ = ). In fact, in the magnetic field parallel to 3C axis the ground non-Kramers doublet is split into two Zeeman sublevels with the energies

( ) ( )ˆ ˆ1, 1 1, 1L S B e Z Z Z L S B e ZM M g S κ L M M g κµ µ= ± = − = ± = = + H H . (4)

On the contrary, this level is non-magnetic in a perpendicular magnetic field. This means that providing 0∆ < the Mn(III) ion possesses uniaxial magnetic anisotropy with

||χ χ⊥> which corresponds to the existence of the easy axis of magnetization. This negative single-ion magnetic anisotropy leads to the negative uniaxial magnetic anisotropy of the entire cluster and to the formation of the barrier for the reversal of magnetization.

The exchange coupling for each such Mn(III) - Mn(II) pair can be roughly described by the isotropic Heisenberg-Dirac-Van-Vleck (HDVV) Hamiltonian

( ) ( )ˆ ˆˆ 2 Mn III Mn IIEXH J= − S S (5) Here we consider a simplifying assumption about isotropic HDVV form of the exchange

Hamiltonian which was proposed by Lines in his consideration of the Co(II) clusters [39]. If the exchange coupling is weak as compared with the SOC splitting, the basis can be truncated to the manifold ( ) ( ) ( ) ( )IIMnIIIMn 2511 =⊗=±= SM,M LL

involving only the ground non-Kramers doublet as illustrated by Fig. 5. The initial isotropic exchange Hamiltonian of HDVV form, Eq. (5), projected onto the ground manifold of the pair, is reduced to the following axi-ally anisotropic effective Hamiltonian of the Ising form:

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( ) ( )ˆ ˆˆ 2 Mn III Mn IIeff Z ZH J S S= − . (6) The effective exchange Hamiltonian for the entire Mn5-cyanide cluster is given by:

( ) ( ) ( ) ( ) ( ) ( )5ˆ ˆ ˆ ˆ ˆˆ Mn 2 1 2 3 4 5eff Z Z Z Z ZH J S + S S S S = − + + (7)

Its eigenvalues are the following: ( ) ( ) ( ) ( )1, 2 , 1, 2,3 2 1, 2 1, 2,3 ,S S S SE M M J M M= − (8)

The microscopic magnetization is equal to ( ) ( ) ( )1,2 3,4,5B e S e Sg M g Mµ κ+ + . Figure 6 shows the energy pattern calculated for J < 0 (Fig. 6a) and J > 0 (Fig. 6b). The

Fig. 5. Energy scheme illustrating the effective exchange interaction acting within the ( ) ( )Mn III

1, 1L LM M= ± = ( ) ( )Mn II5 2S⊗ = - manifold of the Mn(III) - Mn(II) pair.

main features of both energy patterns in Fig. 6 is that in the low-lying groups of levels the energy is decreased with the increase of Zµ and so these groups of levels (shown by red in Fig. 6) form the barriers for the reversal of magnetization. Such energy patterns correspond to the existence of easy axis of magnetization for which ||χ χ⊥> as in the case of pure spin SMMs with 0SD < . The low-lying levels are equidistant and in this aspect the present energy

patterns are drastically different from those described by the ZFS spin Hamiltonian 2ZS SD . In

the last case the energy gaps between the levels spaced near the top of the barrier are smaller than those for levels situated near the bottom of the barrier. This difference arises from the fact that the ZFS describes the second-order anisotropy, while the magnetic anisotropy de-scribed by Eq. (8) is the result of the unquenched orbital angular momentum. Being the first-order effect, the magnetic anisotropy associated with the un quenched orbital angular momen-tum is expected to be much stronger than the second-order anisotropy in spin clusters, and the corresponding barrier for magnetization reversal can be also much larger than the barrier in spin-clusters.

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Fig. 6. Energy patterns of Mn5-cyanide cluster calculated with the aid of Eq. (8) providing antiferromagnetic (left) and ferromagnetic (right) exchange coupling. The low-lying levels forming the barriers are shown in red.

The analysis of the magnetic behaviour of the Mn5-cyanide cluster [23] shows that the real situation in this system is far from the above described idealized picture corresponding to the strong trigonal CF limit. Indeed, the best fit value 1| | 251cm−∆ = of the negative trigonal CF parameter only slightly exceeds the found value 1| | 144 cmκ λ −≈ of the splitting caused by the SOC. Nevertheless, even under these much less favourable conditions the low-lying energy levels calculated for this cluster with the best-fit parameters 1251cm−∆ = − and 1cm83 −−= .J were shown to form the magnetization reversal barrier [24]. This means that the conclusion about the appearance of the barrier in the limit of strong negative trigonal CF proves to be also valid for the arbitrary negative value of ∆ that is compatible with the observed SMM behaviour of the Mn5-syanide cluster.

1.2 Magnetization reversal barrier in cyano-bridged 3d-4d-3d single molecule magnets Recently new [ ( )( ) ( )

5N Me 2 272Mn L H O Mo CN 6H O ⋅

cyanide compound exhibiting dis-

tinct SMM behaviour has been reported [36]. This Mn2Mo-cyanide compound is based on the central pentagonal bipyramidal [MoIII(CN)7]4- heptacyanometalate of D5h – symmetry coupled with two terminal Mn(II) ions via superexchange (Fig. 7a). The effective barrier

1eff 40.5 cmU −≈ is the record for cyanide-bridged SMMs, and the relaxation time at 1.8 K has

been found to be τ ≈ 2.73 × 106 s ≈1 month. In the CF of D5h symmetry the one-electron 4d-level is split into the ground orbital dou-

blet (dZX, dZY) [40] excited doublet (dXY, dX2

-Y2) and the upper singlet (dZ

2). The gap between the two low lying doublets is found to be around 30000 cm−1. Since the CF induced by the carbon atoms is rather strong, the three 4d-electrons of the Mo ion occupy the lowest orbitals dZX and dZY .This gives rise to the ground low-spin orbital doublet 2

1E′′ , which means that along with spin S = 1/2 the angular momentum projection ML = ±1 can be attributed to this term (Fig. 7b).

The SOC acting within this state has the axial form ˆˆZ ZL Sκ λ− and splits the

1, 1 2LM S= ± = level into two Kramers doublets as shown in Fig. 8a. The SOC parameter for the ground orbital doublet is negative, and hence the ground Kramers doublet possesses

1, 1 2L SM M= ± = , while for the excited doublet

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Fig. 7. Molecular structure of the Mn2Mo cluster (a) and electronic configuration and ground term of the Mo(III) ion in [MoIII(CN)7]4- heptacyanometalate (b).

1, 1 2L SM M= ± = ± . The gap | |κ λ ranges from 600 to 1000 cm-1 [41] which is significantly smaller than the CF gap between the ground and excited CF terms. Let us consider now the exchange-coupled Mo(III)-Mn(II) pair. As in the case of Mn5-cyanide cluster we will assume that the initial exchange Hamiltonian is of

Fig. 8. SOC splitting of the 21E′′ - term of the Mo(III) ion, and effective exchange interaction

acting within the ground ( ) ( )Mo III1, 1 2L LM M= ± = ( ) ( )Mn II

5 2S⊗ = - manifold of the

Mo(III)- Mn(II) pair.

isotropic HDVV form. Projecting the HDVV exchange Hamiltonian onto the truncated ground ( ) ( ) ( ) ( )IIMnIIIMo 25211 =⊗=±= SM,M LL - manifold of the Mo-Mn pair (Fig. 8b) we obtain axially anisotropic Ising Hamiltonian of the same form as the Hamiltonian, Eq. (6), found for the Mn(III)-Mn(II) – pair. Then the effective exchange Hamiltonian for the entire Mn2Mo-cyanide cluster is the following:

( ) ( ) ( )2 2ˆ ˆˆ Mn Mo 2 Mo Mneff Z ZH J S S= − (9)

where ( )2ˆ MnZS is the Z-component of the spin operator for the pair of the terminal Mn(II)-

ions. This gives the following set of eigenvalues: ( ) ( ) ( ) ( )2 2Mo , Mn 2 Mo Mn ,S S S SE M M J M M= − (10)

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where ( )Mo 1 2SM = ± and ( )2MnSM takes on the values 0, 1± , 2± , 3± and 5± . Each level in Eq. (10) can be characterized by the value of the total angular momentum projection

( ) ( ) ( ) ( ) ( )2 2Mo Mo Mn Mn MoJ S L S S SM M M M M M= + + = − .

Fig. 9. Energy patterns of Mn2Mo-cyanide cluster. The low-lying levels forming the barrier are shown in red.

Alternatively, each level can be characterized by the microscopic magnetization Zµ for which we find ( ) ( ) ( ) ( ) ( )2 2Mo Mo Mn 2 2 Mo MnZ B e S L e S B S Sg M M g M M Mµ µ κ µ= − + = + providing 2eg = and 1.κ = The equidistant low lying levels form the barrier for magnetiza-tion reversal whose height is determined by the magnitude of the exchange parameter (levels shown in red in Fig. 9). The two examples so far considered show that strong negative single-ion magnetic anisotropy associated with the first-order orbital angular momentum can give rise to the ap-pearance of considerable magnetization reversal barrier composed of equidistant energy lev-els. This barrier is maximal in the strong axial CF limit and determined by the strength and the sign of the exchange coupling.

1.3 Remarks about the role of orbitally-dependent exchange

In general, the HDVV Hamiltonian is applicable when the ground levels of the coupled ions are orbitally non-degenerate, for instance, in the case of high-spin d5 ions in octahedral CF ( ( )6 3 2

1 2g g gA t e ‒ term) or, half-filled subshell ( ( )4 32 2g gA t ). Alternatively, the HDVV Hamil-

tonian is valid when the low-symmetry CF removes the degeneracy and in this way stabilizes orbital singlet. When the orbitally degenerate terms of the constituent ions are involved, the HDVV Hamiltonian, is, in general, inapplicable and a more general, so-called orbitally de-pendent effective Hamiltonian should be employed (see detailed discussion in Ref. [42]). Keeping this in mind, we note here that some cases (indicated in ref. [42]) may prove the ex-ception. In particular, the HDVV Hamiltonian can serve as a good approximation when the kinetic exchange interaction arises mainly from electron transfer within orbitally-nondegenerate sub-shells of the orbitally-degenerate metal ions. This approximation also works well when the number of equal hopping parameters contributing to the kinetic ex-change proves to be equal to the total orbital multiplicity of the system. In such cases we ar-rive at the Ising-type effective exchange Hamiltonian by projecting the HDVV exchange Hamiltonian onto the ground manifold of the pair involving the ground non-Kramers (case of the Mn5-cyanide cluster) or Kramers (case of the Mn2Mo-cyanide cluster) doublet of the or-bitally-degenerate ion.

Turning back to the case of the Mn2Mo-cyanide cluster, we can see that for strictly line-ar Mo-CN-Mn bridging group we have two equal transfer integrals of π-type

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( ) ( ), ,XZ XZ YZ YZt d d t d d= so that this number of transfer parameters is just equal to the total orbital multiplicity of the system (Fig. 10). As to the σ-transfer it is seen from Fig. 10 that it involves only the orbitally nondegenerate sub-shell of the Mo(III) ion and so this transfer also results in the HDVV - type contribution. Therefore, for the system with linear Mo-CN-Mn bridging group the kinetic exchange Hamiltonian is given by:

( ) ( )ˆ ˆˆ 2 Mo III Mn II .EXH J= − S S (11) As was mentioned such HDVV- type Hamiltonian is reduced to the effective Hamilto-

nian of the Ising form, Eq. (9), that is responsible for the formation of the magnetization re-versal barrier as shown in Fig. 9. In Ref. [43] such Ising-type effective Hamiltonian was de-rived for the Mo-Mn pair directed along the C5 axis of the MoIII(CN)7 by considering the kinetic exchange constrained within the ground Kramers-doublet space. In a more general sense, the procedure includes the derivation of the Hamiltonian acting in the full space with subsequent projecting of this Hamiltonian onto the ground manifold as described here. Both these approaches are equivalent and lead to the Ising-type effective Hamiltonian for the linear system. In more general situations the HDVV model proves to be inapplicable and the kinetic exchange becomes essentially orbitally dependent. Actually, this means that the coupling be-tween ions cannot be expressed in terms of spin variables only. The exchange interaction be-tween orbitally degenerate ions includes three different kinds of contributions: spin-spin term (resembling the HDVV Hamiltonian), orbital-orbital coupling involving matrices acting in orbital spaces and spin-orbital terms including products of orbital matrices of one of the ion and the spin operators of other ion (see detailed discussion in Ref. [42]). It is appropriate to emphasize that it is not only a formal mathematical complication of the theory of the magnetic exchange. The most significant observation is that the orbitally-dependent parts of the ex-change interaction give rise to a strong magnetic anisotropy which is an inherent property of the systems comprising orbitally degenerate ions. In this sense, it is worth to mention again that the HDVV Hamiltonian is fully isotropic. The most general form of the orbitally-dependent exchange Hamiltonian for the A-B pair of metal ions, in which the ion A possesses unquenched orbital angular momentum while another ion B has only spin, is the following:

( ) ( )1 2ˆ ˆˆ ˆ ˆ + ,orb dep orb orb

ex A BH R A R A− = S S (12)

where ( )1ˆ orbR A and ( )2

ˆ orbR A are the orbital operators (combination of orbital matrices) for the site A whose explicit forms depend on the overall symmetry of the pair (determining the al-lowed transfer pathways), the electronic configurations of the orbitally-degenerate ions and the symmetry of the local CFs acting on these ions. Upon projecting this Hamiltonian onto the ground manifold we obtain the effective Hamiltonian whose most general form is the follow-ing:

( ) ( ) ( ) ( ) ( ) ( )ˆ ˆ ˆ ˆ ˆ ˆˆ 2 2 2 .eff ZZ Z Z XX X X YY Y YH J S A S B J S A S B J S A S B= − − − (13)

Providing linear geometry of the A-bridge-B group we have two different parameters || , ZZ XX YYJ J J J J⊥≡ = = and so the effective Hamiltonian is axially symmetric:

( ) ( ) ( ) ( ) ( ) ( )||ˆ ˆ ˆ ˆ ˆ ˆˆ 2 2 .eff Z Z X X Y YH J S A S B J S A S B S A S B⊥

= − − + (14)

This form of the Hamiltonian is compatible with the axial point symmetry of the pair, while the values of the exchange parameters depend on the nature of electronic configurations and electronic terms of the involved ions. In the case of a bent

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Fig. 10. Kinetic mechanism of Mo(III)-CN-Mn(II) of the HDVV superexchange in the case of linear geometry.

geometry of the A-bridge-B group the Hamiltonian is, in general, triaxial and has the form of Eq. (13). This situation is apparently unfavourable for the creation of magnetization reversal barrier (especially if the difference between XXJ and YYJ parameters is large) because it pro-motes a fast quantum tunnelling of magnetization effectively decreasing the barrier. Fortu-nately, in some important special cases the actual symmetry of the kinetic exchange Hamilto-nian proves to be higher than the point symmetry of the A-bridge-B group. This occurs, for example, in the case of Mn2Mo-cyanide cluster which is shown in Fig. 7a. For this system, in spite of bent geometry of the bridging Mo(III)-CN-Mn(II) - group, the effective Hamiltonian was found to be of axial form, Eq. (14), with ||| | | |J J⊥>> [37] (close to the Ising limit). In this case MJ remains good quantum number and the low-lying levels form the barrier which however has irregular structure unlike the barrier shown in Fig. 9. Less favourable situation when the bent geometry leads to the triaxial anisotropy is exemplified by the Os(III)-CN-Mn(II) superexchange in trimeric cluster (NEt4)[Mn2(5-Brsalen)2(MeOH)2Os(CN)6] (Fig. 11a) that was shown to exhibit SMM behavior with Ueff ≈ 13 cm-1 [32]. For hypothetical linear system shown in Fig. 11b both ( )orb

1F A and ( )orb2F A operators are proportional to provided

that only the π-transfer contributes to the kinetic exchange. The SOC splits the low-spin ( )2 5

2 2g gT t -term of the Os(III) ion in a perfect octahedral CF (Fig. 11c) into the Kramers dou-blets with J = 1/2 (ground doublet) and J = 3/2 (excited doublet) as shown in Fig. 11d, where J is the quantum number of the total angular momentum of the Os ion. Projecting the terms ∼ ( )2ˆ OsZL and ( ) ( ) ( )2 ˆ ˆˆ Os Os MnZ BL S S on the ground ( ) ( ) ( ) ( )Os III Mn III

1 2 2J S= ⊗ = - manifold we

arrive at the Ising-type effective Hamiltonian. Unfortunately, the real geometry of the cluster exhibits bent geometry (Fig. 11a) and such bending opens additional transfer pathways giving rise to the triaxial effective exchange Hamiltonian [34]. This is the reason why in spite of stronger 5d-3d superexchange, as compared with the 4d-3d one, the Mn(III)2Os(III)-cyanide cluster exhibits lower magnetization reversal barrier as compared with the Mn(II)2Mo(III)-cyanide cluster.

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Fig.11. Molecular structure of the Mn(III)2Os(III)-cyanide cluster (a), linear Mn(III)2Os(III) –

trimer (b), electronic configuration of the Os(III) ion in strong octahedral ligand field of carbon atoms (c) and spin-orbital splitting (d).

One should also mention that in some cases, in spite of Ising-type effective exchange interaction, the magnetization reversal barrier does not appear. Thus, the trigonal bipyramidal cyanide-bridged NiII

3OsIII2 cluster with linear bridging groups Os(III)-CN-Ni(II) does not be-

have as SMM due to the fact that anisotropy axes associated with different Os(III)-Ni(II) pairs, exhibiting Ising exchange, are non-collinear and hence the overall anisotropy of the cluster is not of the Ising type [44].

2. SINGLE-ION MAGNETS OF TRANSITION METAL AND LANTHANIDE IONS WITH UNQUENCHED ORBITAL ANGULAR MOMENTA

The use of highly anisotropic orbitally degenerate ions as building blocks allows not on-ly to create SMMs based on comparatively small clusters but even to obtain SMM behavior for mononuclear complexes. Such behavior was first discovered in lanthanides [45-56], but more recently many mononuclear complexes of nd-ions exhibiting SMM properties have been synthesized and magnetically characterized [57-68]. Hereunder we will briefly discuss several examples of SMMs based on mononuclear complexes which are often called single ion magnets (SIMs).

2.1 Magnetization reversal barrier for two-coordinate Fe(I)-complex: role of low coordination number

Recently the remarkable two-coordinate Fe(I)-complex [Fe(C(SiMe3)3)2]−(Fig. 12a) have reported in Ref. [58]. This system exhibits slow magnetic relaxation below 29K in the absence of the direct current (dc) field, a large effective spin-reversal barrier Ueff ≈ 226 cm-1

and magnetic blocking below 4.5 K. Based on the ab initio calculations [58] the three main factors responsible for the pronounced SMM behavior of this complex have been mentioned: the axial CF splitting imposed by the low coordination number, the SOC of the ground CF multiplet, and the advantage provided by the properties of the basis states with respect to the time reversal (Kramers theorem). The last hinders quantum tunnelling of magnetization that becomes possible only if the hyperfine coupling with the nuclear spin of iron ion is included.

The axial CF leads to the orbital scheme depicted in Fig. 12b [58]. The distribution of seven electrons over these orbitals gives rise to the ground orbital doublet 4E (3d7) whose components possess the same symmetry properties as the 2 2X Y

d−

and XYd orbitals. Since the

linear combinations of these orbitals ( )2 23 2 2d XYX Yd i dϕ ± −

= ± are characterized by the an-

gular momentum projection 2lm = ± the term 4E can be regarded as the state with 2LM = ± . This term undergoes

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Fig. 12. Molecular structure of the complex [Fe(C(SiMe3)3)2]−(a), electronic configuration and the ground 4E -term of this complex (b), and spin-orbital splitting of the 4E –term (c).

a further splitting caused by the SOC which is described by the axial operator acting within the 4E– term:

( )4 ˆˆ ˆSO Z ZH E L Sκ λ= , (15)

where ( ) 12 120 cmSλ ζ −= − ≈ − . The SOC splits the 4E– term into four equidistant

Fig. 13. Energy pattern of the complex [Fe(C(SiMe3)3)2]− calculated with 1κ = . The low-lying levels forming the barrier are shown in red.

Kramers doublets characterized by the absolute value of the projection MJ of the total angular momentum as shown in Fig. 12c. The energy plotted as function of the expectation value of the operator ( )ˆˆˆZ B Z e ZL g Sµ µ κ= + with 1κ = and 2eg = is shown in Fig. 13. It is seen

that the three lowest doublets can be regarded as the magnetization reversal barrier. How-ever, the experimentally observed value for Ueff proves to be smaller than the height of the barrier in Fig. 13 and it is quite close to the energy gap 2 | |λ between the ground doublet with |MJ| = 7/2 and the first excite doublet with |MJ|= 5/2. This is probably due to the phonon assisted quantum tunnelling between the components MJ = −5/2 and MJ = 5/2 of the first ex-cited doublet as schematically shown in Fig. 13. Anyway the found value of Ueff is the largest one observed so far for the SMMs based on 3d – ions, thus showing that low coordination number plays a very constructive role in the design of SIMs because it increases the axial magnetic anisotropy giving rise to the high magnetization reversal barrier.

( )Z Bµ µ

≈ Ueff

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2.2 Single-ion magnets of 4f-ions with unquenched angular momenta In spite of the large number of nd-type SIMs reported to date, the SIMs and SMMs

based on complexes of 4f-ions with unquenched orbital angular momenta seem to be more promising for practical applications. There are two reasons for that. First, 4f-complexes can exhibit (under appropriate symmetry conditions) much stronger magnetic anisotropy [53] than d-complexes. The second reason relates to the much weaker spin-phonon interactions in 4f-complexes because 4f - shell is shielded and 4f-orbitals are less extended. As a consequence, 4f-complexes can show much slower spin-phonon relaxation as compared with the nd-complexes.

The SIM behaviour of lanthanide complexes was first discovered [45,46] for two phthalocyanine double-decker complexes [Pc2Ln]– TBA+ (Ln = Tb or Dy; TBA+ = tetrabu-tylammonium cation) (Fig. 14a). They were found to exhibit temperature and frequency de-pendences of the alternating current (ac) magnetic susceptibility similar to those earlier ob-served for the transition metal based SMMs. These compounds were the first lanthanide metal complexes functioning as SMMs. Moreover, they were the first mononuclear metal complex-es showing SMM properties. One of the most remarkable features of these mononuclear lan-thanide SMMs is that they show slow relaxation of magnetization in temperature ranges that are significantly higher than those for the discovered transition-metal SMMs. Thus, the [Pc2Tb] and [Pc2Dy] complexes exhibit out-of-phase ac susceptibility Mχ′′ peaks at 40 and 10 K with a 103 Hz ac field, respectively [45,46]. Later on the Ho complex with the same struc-ture was also shown to exhibit SMM behaviour. Such characteristic features of SMMs as hys-teresis and resonant quantum tunnelling of magnetization were found in the Tb, Dy and Ho complexes [47,48].

As distinguished from nd-ions, in lanthanides the SOC acts as the leading interaction, stabilizing the ground terms 2S+1LJ (4f n) of the free 4f-ions, e. g. 7F6 (4f 8) for Tb(III) and 6H15/2(4f9) for Dy(III). These terms undergo further splitting by the CF giving rise to the ener-gy patterns of the Stark sublevels. Although the CF in lanthanides is much weaker than in nd-complexes, this interaction is of primary importance because it can lead to the appearance of considerable magnetic anisotropy responsible for the formation of the magnetization reversal barriers. The nearest coordination environment of 4f ion in the [Pc2Tb] and [Pc2Dy] complex-es is represented by eight nitrogen atoms forming two parallel plains rotated with respect to each other by the angle ϕ (Fig.14b). In the cases of [Pc2Tb] and [Pc2Dy] complexes the skew angle ϕ is close to 45o and hence, in the octacoordinated lanthanides LnN8 the local sym-metry can be approximately considered as D4d.

In the case of D4d symmetry the effective CF Hamiltonian acting within the ground ( )n

JS fL 412 + term is written in the operator equivalent form as follows:

0 2 0 0 4 0 0 6 02 2 4 4 6 6

ˆ ˆ ˆCFH A r O A r O A r Oα β γ= + + , (16)

where kqk rA are the ligand field parameters, q

kO are the irreducible tensor operators (Ste-

vens operators [23]) defined with respect to the quantization Z axis (C4 axis of the complex), finally βα , and γ are the Stevens coefficients [69,70].

It was proposed in Ref. [45] to search the sets of the CF parameters, which simultane-ously reproduce the temperature dependence of the dc magnetic susceptibility and the para-magnetic shifts of 1H NMR spectra. The patterns of the Stark sublevels calculated in this way for [Pc2Tb] and [Pc2Dy] complexes are shown in Fig. 15. It is seen from Fig. 15a that the lowest doublet of [Pc2Tb] possesses the largest | |JM val-ue ( 6JM = ± ) and the energy gap between this doublet and the first excited doublet with

268

5JM = ± exceeds 400 cm-1. This situation is similar to that occurring in SMMs based on the transition metal spin clusters, but the magnetization reversal barrier separating sublevels

6+=JM and 6−=JM proves to be much higher as compared to the barriers in spin clusters, 1230 cmeffU −≈ for the [Pc2Tb] complex. The

Fig. 14. (a) Molecular structure of the anion of bis(phthalocyaninato)-lanthanide, and (b) Ln(III) ion surrounded by eight nitrogen atoms showing approximate D4d symmetry of the complex at skew angle ϕ = 45o.

Stark sublevels for the DyIII ion are distributed more evenly, so the SIM behaviour of [Pc2Dy] complex is less pronounced than that for [Pc2Tb]. Indeed, it follows from Arrhenius analysis that 1230 cmeffU −≈ for the [Pc2Tb] complex, while for [Pc2Dy] complex much smaller barrier

of 128 cmeffU −≈ was found. In addition to the complexes of Tb and Dy, the [Pc2Ho] complex was also shown to exhibit SIM properties.

Fig. 15. Stark sublevels of [Pc2Tb]– TBA+(a) and [Pc2Dy]– TBA+(b) calculated with the pa-rameters from Ref. [45]. The levels forming the barriers are red.

Later on it was demonstrated that the concept of SIMs can be extended to other families of mononuclear 4f - complexes. Thus, the polyoxometalate complexes encapsulating some of 4f – ions were shown to exhibit SIM behaviour for coordination sites close to the antipris-matic D4d symmetry [49,50]. Here we will mention one such family, namely [Ln(W5O18)2]9- (LnIII = Tb, Dy, Ho, and Er) whose structure is shown in Fig. 16a. Among these complexes only Er shows slow relaxation of magnetization above 2 K, while the Tb, Dy and Ho com-plexes do not exhibit distinct SIM behavior at T > 2K. Thus the pattern of the Stark sublevels

-10 -8 -6 -4 -2 0 2 4 6 8 10

0

200

400

600

MJ = − 6 MJ = +6

− 5 + 50 + 1- 1

- 2 + 2+ 4- 4 - 3 + 3

( )Z Bµ µ

Ener

gy(c

m-1

)

-12 -10 -8 -6 -4 -2 0 2 4 6 8 10 12

0

200

400

600

MJ = +13/2MJ = -13/2

+11/2-11/2

+9/2-9/2

+7/2-7/2

+5/2-5/2

-3/2 +3/2+1/2-1/2

+15/2-15/2

( )Z Bµ µ

Ener

gy(c

m-1

)

269

calculated with the parameters 0 2 12 36.8 cm ,A r −= − 0 4 1

4 89 cm ,A r −= − and 0 6 16 5.2 cmA r −= −

found by fitting the dc magnetic properties of the [TbW10O36]9- complex [50] shows no mag-netization reversal barrier (Fig. 16b) as distinguished from the energy levels of the [Pc2Tb] complex which were shown to form the barrier (Fig. 15a). On the other hand, in contrast to the system [ErW10O36]9-, the [Pc2Er] complex does not exhibits SIM properties. The difference between the magnetic properties of these two classes of complexes be-longing to the same symmetry lies in the fact that the nearest ligand surrounding of the lan-thanide encapsulated in polyoxomatalate represents axially compressed square antiprism, meanwhile in phthalocyaninato complexes such antiprism is axially elongated. One can thus conclude that the axially elongated sites promote the SIM behavior in cases of Tb and Dy complexes, as exemplified by the double-decker bis-(phthalocyaninato) complexes, while axially compressed sites in the [Ln(W5O18)2]9- complexes are favorable to obtain Er(III)-based SIM. This can be realized by considering the distribution of the point charges around the 4f-ion [71]. In the framework of the point charge model one can use the following expressions for the CF parameters:

( ) ( )( )

2800

0 11

4 1 ,2 1

p i p i ipp p p

i i

Z e YA r c

p R

p σ θ ϕ+

=

−=

+ ∑ , (18)

where iiR θ, and iϕ are the polar coordinates of the i th ligand of the nearest surrounding of the lanthanide ion, eZi− is the charge of the i th ligand, ( )iipY ϕθ ,0 are the spherical

harmonics, pσ are the shielding parameters, finally 0pc (p=2, 4, 6) are the following numer-ical factors. For the sake of brevity we do not give these parameters and CF parameters here. The energies of the D4d complex in the framework of the point charge CF are fully de-termined by the two structural factors, namely, by the distance R between the Ln(III) ion and the atoms of nearest ligand surrounding and on the polar angle θ . Alternatively, one can characterize the geometry of the complex by other two values, namely, by in-plane distance

ind , and the interplane distance ppd (see Fig. 17). Simple geometrical consideration gives the following dependence between these two sets of values: ( ) .

2cos ,2

2222

21

inpp

ppinpp

dd

dddR

+=+= θ (19)

Providing undistorted square in antiprismatic geometry ( )( , cos 1 3in ppd d θ= = ) the pa-

rameter 0 22 0A r = . For axially elongated sites ( )( , cos 1 3in ppd d θ< > ) we find that

0 22 0A r > and, hence, the term 0 2 0

2 2ˆA r Oα tends to stabilize the doublet with large |MJ|

value for lanthanides with negative Stevens coefficient α (i. e. for Tb and Dy, see Table 1) and the doublet with small |MJ| in case of ions with positive α (i. e. for Er ion). As the rule, the 0 2 0

2 2ˆA r Oα contribution dominates except for the case of in ppd d= when 0 2

2A r is small

and the energy pattern is mainly determined by the contribution 0 4 04 4

ˆA r Oβ . Thus, we arrive at the conclusion that axial elongation is favourable for SIM behaviour of Tb and Dy com-plexes in agreement with what

270

Fig. 16. Polyhedral view of [TbW10O36]9- polyoxoanion (a), and of the Stark sublevels calcu-lated for this complex with the parameters from Ref.[50].

is observed for double-decker bis-(phthalocyaninato) complexes. In contrast, for axially com-pressed sites ( ), cos 1 3in ppd d θ> < , 0 2

2 0A r < and hence the SIM behaviour is expecta-ble for ions with positive Stevens coefficient α (i. e. for the Er ion, Table 1). This takes place, for example, in the [Er(W5O18)2]9-complex. Note that the parameters 0 2

2 ,A r 0 44A r

and 0 66A r do not depend on the skew angle ϕ and so the above arguments are also valid for

the case of C4 symmetry when ϕ ≠ 45o (and

Fig. 17. The parameters determining the geometrical structure of the D4d complexes.

ϕ ≠ 0, 90o) and also for the case of cubic Oh symmetry when ϕ = 0 or 90o. Note that in the latter case 0 2

2 0A r = and so the cubic geometry (Oh symmetry) is less suitable for obtaining

SIMs than the geometry of antiprism (D4d symmetry). It is also notable that at ϕ ≠ 45o the CF Hamiltonian includes along with the terms 0 2 0

2 2ˆA r Oα , 0 4 0

4 4ˆA r Oβ

and 0 6 06 6

ˆA r Oβ also

two additional off-diagonal terms 4 4 44 4

ˆA r Oβ and 4 6 46 6

ˆA r Oγ which mix the MJ states

with 4=JM∆ promoting thus quantum tunnelling of magnetization. The analysis based on

the point charge model shows that the parameter 4 44A r is vanishing only providing ϕ =45o

and reaches the maximal value for cubic geometry (ϕ =0 or 90o), while the parameter 4 66A r

= 0 for D4d and Oh symmetries. The application of the point charge model for similar analysis of other typical environ-

ments, such as triangular dodecahedron and trigonal prism, has allowed to establish the fol-

271

lowing simple rules which are advantageous for the rational design of SIMs based on 4f-complexes [71]: 1. As a general rule, to form an energy barrier leading to slow spin relaxation the pattern of the Stark sublevels should exhibit ground state with high |MJ | and small mixing between the +MJ and −MJ components in order to reduce the relaxation through quantum tunnelling of magnetization. The most trivial condition to get a high-MJ ground-state doublet is to have a large J values which occur for the second half of the lanthanide series, with the Tb(III), Dy(III), Ho(III), Er(III) and Tm(III) complexes being the best choices. 2. The second condition for having high | MJ | ground state and the magnetization reversal barrier is the high symmetry of the complex. This can be achieved for complexes with a pseu-do-axial symmetry like D4d, C5h, D6d, or for any symmetry of order 7 or higher. The most suit-able situation occurs when the second-order uniaxial anisotropy determined by the parameter

0 22A r α

(also known as D) dominates. For example, comparing two high-symmetric octa-

coordinated complexes, one with the antiprismatic D4d symmetry and another one with the cubic symmetry ( 0 2

2 0A r = ), we could see that the Oh geometry is unfavorable to exhibit a large barrier for the magnetization reversal. In contrast, the systems with D4d symmetry have strong uniaxial anisotropy (either positive or negative), which is provided by many SIMs. 3. In most cases the off-diagonal terms 4 4 4

4 4ˆA r Oβ , 4 6 4

6 6ˆA r Oγ

etc. allows quantum tun-nelling of magnetization and they also effectively reduce the barrier. In some cases, however, such terms cannot lead to the quantum tunnelling in the ground state and thus they do not pre-clude from SIM behavior of the complex. Such situation is exemplified by the Ho(III) com-plex of D2d symmetry. The calculated splitting diagram for the J = 8 ground state of Ho(III) in this environment evidences that the two components of the ground-state doublet are com-posed by the following M J values: (+7, +3, − 1, − 5) and ( − 7, − 3, +1,+5). It is seen that alt-hough these two functions are formed by an extensive mixture of different M J states, they cannot participate in tunnelling because there is no overlap between these two sets of M J val-ues. Another way to avoid the destructive effect of quantum tunnelling is to use lanthanide ions with non-integer J – values (e. g. Dy(III) or Er(III) ions). In this case the quantum tunnel-ling is forbidden by the Kramers theorem and becomes allowed (but rather weak) only through the hyperfine coupling of the 4f – electrons with nuclei possessing non-integer spin values, so that the total electron-nuclear angular momentum of the complex becomes integer.

In summary, one can say that the main requirement to obtain good SIMs based on lan-thanide complexes is to have strongly axial CF. In view of this it has been recently suggested to consider Dy complexes with low coordination numbers (1 and 2) as good candidates to obtain SIMs with unprecedentedly high barriers (see examples in ref. [72] ), but such com-plexes are often not too stable. Fortunately, there is a possibility to obtain almost perfect axial CF with coordination number that is higher than one or two, when such CF is created by two negatively charged axial ligands, while the remaining equatorial ligands are nearly neutral. This has recently led to a series of dysprosium complexes [Dy(OPCy3)2(H2O)5]3+ (Cy = cy-clohexyl) [73], [Dy(OPtBu(NHiPr2)2)2(H2O)5]3+ [74], [Dy(BIPM)2]−, (BIPM ={C(PPh2NSiMe3)2}2−) [75] and [Dy(bbpen)Br] (bbpen =N,N'-bis(2-hydroxybenzyl)- N,N '-bis(2-methylpyridyl)ethylenediamine) [76], for which very high effective barriers of Ueff = 543, 735, 813 and 1025 K, respectively, have been found. Finally, quite recently the complex [Dy(OtBu)2(py)5][BPh4] exhibiting the record barrier Ueff ≈ 1755 K (1220 cm-1) has been re-ported [77]. This is a pentagonal bipyramidal complex of the type [DyX2L5]+, where L are the neutral and X are the anionic donors which induce almost perfect axial CF.

272

2.3 Field induced single-ion magnets with positive axial and strong rhombic anisotropy

In all examples so far discussed the presence of a considerable magnetization reversal barrier formed by uniaxial easy-axis anisotropy represents the key prerequisite for the crea-tion of SMM. Recently a slow magnetic relaxation has been also discovered in some com-plexes of Kramers ions with dominant easy-plane magnetic anisotropy and strong rhombic anisotropy. Such unusual behavior takes place only in the presence of external magnetic dc field and so these complexes are often termed “field induced SIMs”. The majority of field induced SIMs represent the high-spin Co(II) complexes [61-68]. In this Section we briefly discuss one such example representing the recently reported Et4N[CoII(hfac)3] (hfac =hexafluoroacetylacetonate) complex [68], whose structure is shown in Fig. 18. The nearest ligand surrounding of the Co(II) ion formed by six oxygen atoms exhibits strong axial and rhombic distortions.

Fig. 18. Molecular structure of Et4N[CoII(hfac)3] (C – gray, F – light green) and frequency dependences of the out of phase ac susceptibility measured at dc field B = 0.1 T and tempera-tures ranging from 1.8 to 3.3 K with increment of 0.1 K.

Dynamic ac magnetic susceptibility measurements revealed no frequency dependence of the in-phase ( Mχ′ ) and out-of-phase ( Mχ′′ ) signals in the absence of an applied dc field. However, in the presence of a small external field of 0.1T the complex showed typical SMM behaviour (Fig. 18b).

The magnetic properties of the Co(II) complex are often analyzed by the Griffith Hamil-tonian that explicitly takes into account the unquenched orbital angular momentum of the Co(II) ion [78-80]:

( ) ( )2 2 23 1 3ˆ ˆˆ ˆ ˆ ˆ ˆ ˆ12 3 2ax Z rh X Y B eH κλ L L L L L gµ κ = − + ∆ − + + ∆ − + −

LS B S L (20)

This Hamiltonian operates within the basis of the ground octahedral 4T1g – term of the Co(II) ion which represents the mixture of 4T1g – terms arising from the terms 4F (ground) and 4P (excited) of the free Co(II) ion by the cubic component of the CF. The 4T1g – term can be associated with the fictitious orbital angular momentum L = 1 and the spin S = 3/2. Factor 3/2 in SOC operator (first term in Eq. (20)) and in Zeeman operator (last term in Eq. (20)) appears due to the fact that the matrix of the orbital angular momentum operator L defined in the 4T1g(4F) basis differs by this factor from the matrix of L defined in pure atomic 4P – basis. The orbital reduction factor κ describes both the covalence effect and the admixture of the excited 4T1g(4P) term to the ground term 4T1g(4F) by the cubic CF. The second term in Eq. (20)

273

describes the splitting of the 4T1g – term into orbital singlet ( )42 0g LA M = and the orbital dou-

blet ( )4 1g LE M = ± caused by the axial (tetragonal) distortion of the octahedron. Finally, the third term is responsible for the further splitting of the tetragonal orbital doublet induced by the rhombic distortion of the octahedral surrounding of the Co(II) ion.

The sign of the axial CF parameter ax∆ plays crucial role in the magnetic behavior of the Co(II) complex since it determines the sign of the magnetic anisotropy of the system. Thus, providing 0ax∆ < the ground term of the axially distorted complex proves to be an or-bital doublet, and the system exhibits negative magnetic anisotropy || 0χ χ⊥ − < which corre-sponds to the existence of an easy axis of the magnetization. In contrast, for positive ax∆ the anisotropy is also positive which means that the system possesses an easy plane of magnetiza-tion. Note, however, that in the presence of strong rhombic CF the terms “easy axis” and “easy plane” are only of conditional character.

The main problem in fitting the dc magnetic data for the powder sample is that they are only slightly sensitive to the change of the sign of .ax∆ An additional complication arises from the fact that the dc magnetic data prove to be almost independent of rh∆ . As a result, the us-age of the Hamiltonian, Eq. (20), in which ax∆ and rh∆ are regarded as fitting parameters can give a multitude sets of the best fit parameters. Therefore, without additional independent information about the sign of the parameter ax∆ , neither the adequacy of a model nor the cor-rectness of these parameters can be tested. From this point of view the usage in addition to the dc magnetic measurement of complementary spectroscopic techniques, like EPR providing a direct access to the sign of the magnetic anisotropy (incorporated in the principal values of g-tensor for the ground Kramers doublet of the Co(II) ion) has been shown to be quite useful as well as the quantum-chemical evaluation of the parameters ax∆ and rh∆ .

The values 1428.29 cmax−∆ = , 190.34 cmrh

−∆ = obtained through the quantum chemical calculations [68] for the Et4N[CoII(hfac)3] complex indicate the presence of non-uniaxial magnetic anisotropy with strong positive axial and significant rhombic contributions. These values of the CF parameters as well as the free-ion value λ = − 180 cm−1 of the SOC parame-ter for the Co(II) ion have been used in Ref. [68]. The only parameter has been considered as an adjustable one, namely, the orbital reduction factor κ. The best fit value of this parameter is κ = 0.72. It is remarkable that with the only fitting parameter one can reproduce quite well both the dc magnetic data and the effective g-tensor derived from the EPR spectra. Thus, the calculated values 2.51, 4.01, 5.32ZZ XX YYg g g≈ ≈ ≈ of the principal values of g-tensor for the ground Kramers doublet prove to be quite close to those ( 2.502,

ZZg =

4.251, 5.467XX YYg g= = ) obtained from simulation of EPR spectra. The energy levels of the Co(II) ion calculated as functions of the parameter ax∆ for found values 190.34 cmrh

−∆ = , κ = 0.72 and λ = − 180 cm− 1 are shown in Fig. 19. Such kind of plot represents a generalization of the Griffith diagram [78] to the case of tri-axial symmetry. The vertical section marked by dashed red line corresponds to the found axial CF value

1428.29 cmax−∆ = and shows the evaluated energy spectrum consisting of six Kramers dou-

blets. The two low-lying doublets arise from the spin-orbital splitting of the tetragonal 4A2g − term (this splitting can be approximately described by the ZFS spin Hamiltonian 2ˆ

ZD S with positive D value), and the upper four doublets appear as a result of the spin-orbital splitting of the 4Eg − term. The first excited doublet is separated from the ground one by the energy gap of around 178 cm-1, and the second excited doublet lies ≈ 463 cm-1 .

274

Let us briefly discuss the most ambiguous question concerning the origin of slow magnetic relaxation in Kramers ions exhibiting easy plane anisotropy. Several explanations of such slow relaxation have been recently proposed. One explanation is based on the assump-tion that this relaxation is a result of strong rhombic anisotropy,

Fig. 19. Dependences of the energy levels of the Co(II) ion on the parameter ax∆ . The dashed lines restrict the area of ax∆ for which | | 3ax rh∆ < ∆ |.

with the effective barrier for spin reversal being determined by the parameter E through the approximate relation 2 | |effU E≈ . This estimation sometimes gives the values of effU which are close to those found from the Arrhenius plot [61]. For other systems such estimation does not provide correct values of effU and so it has been proposed [81] that relaxation between the levels 1/ 2SM = − and 1/ 2SM = + is slowed by a phonon bottleneck and involves the Orbach process through the first excited doublet with 3 / 2SM = ± . In the framework of this picture the effective barrier is approximated by the energy gap between the ground (±1/2) and excited (±3/2) doublets, that is ( )1 22 22 3effU D E≈ + . In some cases effU found in this way proves to be close to the barrier extracted from the Arrhenius plot but for other systems such explana-tion fails. Thus for the above considered complex Et4N[CoII(hfac)3] the Arrhenius plot shows the presence of the barrier 119.5 cmeffU −≈ [68] that is one order of magnitude smaller than the energy gap (≈178 cm-1) between the ground and first excited Kramers doublets. So one should exclude the Orbach process from the consideration and consider another possibility to reproduce the temperature dependence of the relaxation time by taking into account one-phonon direct processes (dominating in the low temperature region) in combination with the two-phonon Raman processes (important at higher temperatures).

The most comprehensive and non-equivocal explanation of the origin of slow magnetic relaxation in Kramers complexes with an easy plane anisotropy has been given in Ref. [62]. It has been shown that the slow relaxation is a general a consequence of time-reversal sym-metry (van Vleck cancellation) that hinders direct spin–phonon transitions between the sublevels of the ground Kramers doublet. The hyperfine interaction breaks time reversal symmetry and opens channels for direct spin–phonon relaxation. The relaxation between the Kramers-conjugate states can also occur through Orbach and Raman processes albeit, as we have seen, the Orbach process is often irrelevant because of large energy gap between the ground and excited Kramers doublets. At the same time the hyperfine interaction gives rise to a quantum tunnelling that masks the relaxation phenomenon at zero dc field. To summarize, one can say that the three main prerequisites to get slow relaxation in such kind of complexes are [62]: (i) half-integer spin, (ii) strong magnetic anisotropy, (iii) minimized hyperfine inter-action, that is, the usage of Kramers ions having stable isotopes with zero nuclear spin.

275

3. SUMMARY AND CONCLUSIONS In the SMMs reported until now the blocking temperatures do not exceed a few Kelvin,

which are too low for application of these systems as the data-storage units. Therefore, the design of new SMMs with higher blocking temperatures and higher magnetization reversal barriers represents a challenge of molecular magnetism. Here we have briefly discussed a promising approach to increase the magnetic anisotropy by incorporating metal ions with un-quenched orbital angular momenta in clusters.

By considering selected examples of such molecules we have demonstrated that strong first-order magnetic anisotropy arising from SOC,low-symmetry CF and orbitally-dependent superexchange can enhance magnetization reversal barrier. Summarizing the experimental and theoretical results obtained for such systems one can indicate the following general rules to take advantage from the orbital effects and to design highly anisotropic SMMs exhibiting high blocking temperatures: • The nearest ligand surroundings of the constituent metal ions possessing unquenched or-bital angular momenta is to be be organized in such a way, that they produce strong axial CFs, stabilizing the orbital doublets for the d- ions or the ±MJ doublets with large |MJ| for f-ions. These orbital doublets are strongly anisotropic and can give rise to a considerable magnetiza-tion reversal barrier. • As distinguished from the spin-clusters in which the barrier height is determined mainly by the SD value, the barrier height in systems containing orbitally degenerate ions is de-pendent on the strength of the exchange interaction, and so the increase of the exchange cou-pling can be considered as an important ingredient of the design strategy. From this point of view the use of 4d and 5d ions seems to be promising, since they promote a strong exchange coupling due to more extended 4d and 5d – orbitals as compared to the 3d –ones. • All local anisotropy axes in clusters should be possibly collinear in order to increase the global magnetic anisotropy. Several important classes of orbitally-degenerate systems exhibiting slow magnetic re-laxation remained out of the scope of this review. These are, for example, mixed 3d-4f SMMs, polynuclear 4f- complexes, and SIMs based on actinide complexes [51,54,82,83,84,85].

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